DYNAMIC OCEANOGRAPHY

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DYNAMIC OCEANOGRAPHY
JAN ERIK WEBER
Department of Geosciences
Section for Meteorology and Oceanography
University of Oslo
09.11.04
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CONTENTS
1. SHALLOW-WATER THEORY, QUASI-HOMOGENEOUS OCEAN
1.1 Inviscid motion, potential vorticity………………………………………….3
1.2 Linear waves in the absence of rotation……………………………………..8
1.3 Effect of rotation, geostrophic adjustment…………………………………13
1.4 Sverdrup waves, inertial waves and Poincarè waves………………………16
1.5 Kelvin waves at a straight coast……………………………………………23
1.6 Amphidromic systems……………………………………...........................26
1.7 Equatorial Kelvin waves…………………………………………………...29
1.8 Topographically trapped waves……………………………………………31
1.9 Planetary Rossby waves……………………………………………………35
1.10 Topographic Rossby waves………………………………………………..41
1.11 Barotropic instability………………………………………………………42
1.12 Barotropic flow over an ocean ridge………………………………………45
2. WIND-DRIVEN CURRENTS AND OCEAN CIRCULATION
2.1 Equations for the mean motion………………………………......................51
2.2 The Ekman elementary current system……………………………………..57
2.3 The Ekman transport………………………………………………………..61
2.4 Divergent Ekman transport and forced vertical motion(Ekman suction)…..62
2.5 Variable vertical eddy viscosity…………………………………………….65
2.6 Equations for the volume transport…………………………………………68
2.7 The Sverdrup transport……………………………………………………...72
2.8 Theories of Stommel and Munk (western intensification)…………….........75
3. BAROCLINIC MOTION
3.1 Two-layer model………………………………………………………........81
3.2 Continuously stratified fluid………………………………………………..87
3.3 Free internal waves with rotation ………………………………………….88
3.4 Internal response to wind-forcing, upwelling at a straight coast…………...94
3.5 Baroclinic instability………………………………………………............100
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1. SHALLOW-WATER THEORY, QUASI-HOMOGENEOUS
OCEAN
1.1 Inviscid motion, potential vorticity
Governing equations
We study motion in an inviscid fluid with density. The fluid is rotating about the
z-axis with constant angular velocity sin, where  is the latitude of our site of
observation. Furthermore, (x, y) are horizontal coordinate axes along the undisturbed
sea surface, and the z-axis is directed upwards. The position of the free surface is
given by z   ( x, y , t ) , where  is referred to as the surface elevation. The atmospheric
pressure at the surface is denoted by PS(x, y, t). The bottom topography does not vary
with time, and is given by z   H ( x, y ) ; see the sketch in Fig. 1.1.
Fig 1.1. Definition sketch.
The equation of motion for a fluid particle of unit mass can now be written

 

Dv
1
 fk  v   p  gk ,
(1.1.1)
dt


where D/dt  /t + v   is the total derivative following a fluid particle, and f =
2sin is the Coriolis parameter.
We take that the mass of fluid within a geometrically fixed volume only can
change as a result of advection of fluid particles. The basic assumption of
conservation of mass for a fluid can be stated mathematically as
4

D
   v .
dt
(1.1.2)
This is often referred to as the continuity equation. In this section we will assume that
the density of a particle is conserved (D/dt = 0). Thus, the continuity equation
reduces to

v  0.
(1.1.3)
Furthermore we will assume that the density is the same for all particles
(homogeneous fluid). We orientate our coordinate system such that the x-axis is
tangential to a latitudinal circle and the y-axis is pointing northwards; see Fig. 1.2.
Fig. 1.2. Orientation of coordinate axes.
In our reference system f is only a function of y. We may then write approximately
that
 df 
f  f 0    y  f 0   y ,
 dy  0
(1.1.4)
where
f 0  2 sin 0 ,

1 d
2 sin   0
R d


2


cos 0 . 
R

(1.1.5)
This is called the beta-plane approximation. If f is approximately constant in an (x, y)area, we say that the motion occurs on an f-plane.
The kinematic and dynamic boundary conditions at the surface can be written,
respectively, as
5
D
( z   )  0, z   ( x, y , t ) ,
dt
(1.1.6)
p  PS , z   ( x, y , t ) .
(1.1.7)
The kinematic boundary condition at the bottom can be expressed as

v  z  H   0, z   H ( x, y ) .
(1.1.8)
We now integrate the continuity equation (1.1.3) from z = H to z = (x, y, t):



 wz dz    u x dz   v y dz .
H
H
(1.1.9)
H
Utilizing (1.1.6) and (1.1.8), we find that




t    u dz 
v dz .
x  H
y H
(1.1.10)
The basic assumption in shallow water theory is that the pressure distribution in the
vertical direction is hydrostatic. By utilizing that the density is constant, and applying
(1.1.7), this leads to
p  g (  z )  PS ( x, y, t ) .
(1.1.11)
This means, when we return to the vertical component in (1.1.1), that the vertical
acceleration Dw/dt must be so small that it does not noticeably alter the hydrostatic
pressure distribution. We will return to the validity of this in section 1.2. The
horizontal components of (1.1.1) can thus be written
Du
1
 fv   g x  PSx
dt

(1.1.12)
Dv
1
 fu   g y  PSy
dt

(1.1.13)
Throughout this text we will alternate between writing partial derivatives in full, and
(for economic reasons) as subscripts.
We realize that the right-hand sides of (1.1.12) and (1.1.13) are independent of z.
By utilizing that v  Dy/dt and f = f0 + y, (1.1.12) can be written
D
1
1
2
 u  f 0 y   y    g x  PSx
dt 
2


(1.1.14)
From (1.1.14) it follows that D(u  f 0 y   y 2 / 2) / dt is independent of z. Thus, this is
also true for (u  f 0 y   y 2 / 2) , and thereby also for u, if u and v were independent of
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z at time t = 0. Similarly, from (1.1.13) we find that v is independent of z. We can
accordingly write
u  u ( x, y, t ), 

v  v( x, y, t ). 
(1.1.15)
Furthermore, it now follows from (1.1.3) that wz is independent of z. Hence, by
integrating in the vertical:
w  (u x  v y ) z  C ( x, y, t ) .
(1.1.16)
The function C is obtained by applying the boundary condition (1.1.8) at the ocean
bottom. The vertical velocity can thus be written
w  (u x  v y )( z  H )  uH x  vH y .
(1.1.17)
Since u and v are independent of z, the integrations in (1.1.10) are easily performed.
We then obtain the following nonlinear, coupled set of equations for the horizontal
velocity components and the surface elevation
ut  uu x  vuy  fv   g x 
1
PSx ,
(1.1.18)
PSy ,
(1.1.19)
t  u( H   )x  v ( H   )y  0 .
(1.1.20)
vt  uv x  vv y  fu   g y 

1

To solve this set of equations we require three initial conditions, e.g. the distribution
of u, v, and  in space at time t = 0. If the fluid is limited by lateral boundaries (walls),
we must in addition ensure that the solutions satisfy the requirements of no flow
through impermeable walls.
Potential vorticity
We define the vertical component of the relative vorticity in our coordinate
system, e.g. Fig. 1.2, by
  vx  u y .
(1.1.21)
In addition, every particle in this coordinate system possesses a planetary vorticity f,
arising from solid body rotation with angular velocity  sin  . Hence, the absolute
vertical vorticity for a particle becomes f   . We shall derive an equation for the
absolute vorticity. It is obtained by differentiating the equations (1.1.18) and (1.1.19)
by   / y
and  / x , respectively, and then add the resulting equations.
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Mathematically, this means to operate the curl on the vector equation to eliminate the
gradient terms. Since f is independent of time, we find that
D
( f   )  ( f   )( u x  v y ) .
dt
(1.1.22)
By using that H is independent of time (1.1.20) can be written
D
( H   )  ( H   )( u x  v y ) .
dt
(1.1.23)
Here, H +  is the height of a vertical fluid column. We define the potential vorticity
Q by
Q
f 
.
H 
(1.1.24)
By eliminating the horizontal divergence between (1.1.22) and (1.1.23), we find for Q
that
DQ
0.
dt
(1.1.25)
This equation expresses the fact that a given material vertical fluid column always
moves in such a way that its potential vorticity is conserved.
Alternatively, we can apply Kelvin’s circulation theorem for an inviscid fluid to
derive this important result. Kelvin’s theorem states that the circulation of the
absolute velocity around a closed material curve (always consisting of the same fluid
particles) is conserved. For a material curve  in the horizontal plane, Kelvin’s and
Stokes’ theorems yield




v


r

k
abs

  (  vabs )  const . ,

(1.1.26)

where  is the area inside . Furthermore, in the surface integral:


k  (  vabs )  f   .
(1.1.27)
When the surface area  in (1.1.26) approaches zero, we have
( f   )  const.
(1.1.28)
In addition, the mass of a vertical fluid column with base  must be conserved, and
hence
 ( H   )  const.
(1.1.29)
This is valid for all times, since a vertical fluid column will remain vertical; see
(1.1.15). In our case the fluid is homogeneous and incompressible, i.e.  is the same
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for all particles. Thus, by eliminating  between (1.1.28) and (1.1.29), we find as
before that
Q
f 
 const. ,
H 
(1.1.30)
or, equivalently, DQ / dt  0 .
In the ocean we usually have that || << f and || << H. For stationary flow,
assuming that |H| >> || and |f H| >> |H |, (1.1.25) yields approximately that

v  ( f / H )  0 .
(1.1.31)
On an f-plane, this equation reduces to

v  H  0 .
(1.1.32)
Accordingly, the flow in this case follows the lines of constant H (i.e. the bottom
contours). This phenomenon is called topographic steering. On a beta-plane the flow
will follow the contours of the function f/H; see (1.1.31).
Take that the motion is approximately geostrophic. For constant surface pressure,
we then obtain from (1.1.18) and (1.1.19) that
 g 
v  (k   ) .
f
(1.1.33)
For stationary flow the theorem of conservation of potential vorticity reduces

to v  Q  0 . Combined with (1.1.33), this leads to

k  (  Q )  0 .
(1.1.34)
This means that the surfaces of constant  and constant Q coincide. Accordingly, Q
and  are uniquely related, or
   (Q ) .
(1.1.35)
For the special case where we have topographic steering on an f-plane (when Q varies
only with H), we find from (1.1.35) that
   (H ) .
(1.1.36)
Accordingly, for stationary, geostrophic motion with a free surface, the iso-lines for
the surface elevation are parallel to the depth contours.
1.2 Linear waves in the absence of rotation
We assume small disturbances from the state of equilibrium in the ocean, twodimensional motion (/y = 0, v = 0), and constant depth. Furthermore, we take the
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atmospheric pressure to be constant along the surface. The set of equations (1.1.18)(1.1.20) can then be linearized, reducing to
ut   g x , 

t   Hu x . 
(1.2.1)
Eliminating the horizontal velocity, we find
tt  gH xx  0 .
(1.2.2)
This equation is called the wave equation. By inserting a wave solution of the type
  exp( ik ( x  c0t ) , representing a complex Fourier component, we find from (1.2.2)
that c0  (gH )1 / 2 , i.e. all wave components move with the same phase speed
regardless of wavelength. Such waves are called non-dispersive. We need not limit
ourselves to consider one single Fourier component. From (1.2.2) we realize
immediately that a general solution can be written
  F1 ( x  c0t )  F2 ( x  c0t ) .
(1.2.3)
If, at time t = 0, the surface elevation was such that  = F(x), and t = 0, it is easy to
see that the solution becomes

1
F ( x  c0t )  F ( x  c0t ).
2
(1.2.4)
From (1.2.1) and (1.2.4) we find for the acceleration
ut   g x  
g
F ' ( x  c0t )  F ' ( x  c0t ),
2
(1.2.5)
where F ' ( )  dF / d . Hence, the horizontal velocity is given by
u
g
F ( x  c0t )  F ( x  c0t ).
2c0
(1.2.6)
From (1.2.4) we can display the evolution of an initially bell-shaped surface elevation
F(x) with typical width L; see the sketch in Fig. 1.3.
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Fig. 1.3. Evolution of a bell-shaped surface elevation.
We note that the initial elevation splits into two identical pulses moving right and left
with velocity c0 = (gH)1/2. In a deep ocean (H = 4000 m), the phase speed is c0  200
m s1, while in a shallow ocean (H = 100 m) we have c0  30 m s1. If the maximum
initial elevation in this example is h, i.e. F(0) = h, we find from (1.2.6) that the
velocity in the ocean directly below peak of the right-hand pulse can be written
u
gh
,
2c 0
(1.2.7)
when t >> L/c0, that is after the two pulses have split. If we take h = 1 m as a typical
value, the deep ocean example yields u  2.5 cm s1, while for the shallow ocean we
find u  17 cm s1.
As a second example we consider an initial step function:
 1
 2 h, x  0,
F ( x)  
1
 h, x  0.
 2
(1.2.8)
In this case, the velocity and amplitude development becomes as sketched in Fig. 1.4.
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Fig. 1.4. Evolution of a surface step function.
It is obvious that we in an example like this (with a step in the surface at t = 0) must
be careful when using linear theory, which requires small gradients. In a more
realistic example where differences in height occurs, the initial elevation will have a
final (an quite small) gradient around x = 0. Qualitatively, however, the solution
becomes as discussed above.
It is easy to show that the solution for the step problem satisfies the mass and
energy conservation laws: Choose a geometrically fixed area –D  x  D, where D
 c0t. From Fig. 1.4 it is obvious that the mass of fluid within this area is constant.
We choose z = h/2 as the level where the potential energy is zero. The initial
potential energy can thus be written
E p0  MghG 
1
gDh 2 ,
2
(1.2.9)
where M = Dh is the fluid mass above the zero level, and hG = h/2 is the vertical
distance to the centre of mass of this fluid. The initial kinetic energy is zero, since we
start the problem from rest. At a later time t = D/c0, the potential energy becomes
E p  Mg
while the kinetic energy can be written
h 1
 gDh 2 ,
4 4
(1.2.10)
12
Ek
1
1
 (2 DH )u 2  gDh 2 .
2
4
(1.2.11)
Thus, from (1.2.9)-(1.2.11) we find that Ek  E p  E p0 , i.e. the total energy is
conserved. This is of course also true for the bell-shaped elevation in the first
example.
The hydrostatic approximation
Finally, we address the validity of the hydrostatic approximation in the case of
waves in a non-rotating ocean. We rewrite the pressure as a hydrostatic part plus a
deviation:
p  g (  z )  PS  p' ,
(1.2.12)
where p ' is the non-hydrostatic deviation. The vertical component of (1.1.1) becomes
to lowest order:
wt  
1

pz ,
(1.2.13)
while the horizontal component can be written
ut   g x 
1

px .
(1.2.14)
The hydrostatic assumption implies that
1
px  ut .

(1.2.15)
If the typical length scales in the x- and z-directions are L and H, respectively, we
obtain from the continuity equation that
u ~
L
w,
H
(1.2.16)
where ~ means order of magnitude. From (1.2.13) we then find
p'

~
H2
ut .
L
(1.2.17)
Utilizing this result, the condition (1.2.15) reduces to
H 2 / L2  1 .
(1.2.18)
Thus, we realize that the assumption of a hydrostatic pressure distribution in the
vertical requires that the horizontal scale L of the disturbance must be much larger
13
than the ocean depth. For a wave, L is associated with the wavelength; for a single
pulse, L corresponds to the characteristic pulse width.
1.3 Effect of rotation, geostrophic adjustment
We now consider the effect of the earth’s rotation upon wave motion. Linear
theory still applies, and we take the depth and the surface pressure to be constant.
Furthermore, we assume that f is constant. Equations (1.1.18)-(1.1.20) then reduces to
ut  fv   g x ,
(1.3.1)
vt  fu   g y ,
(1.3.2)
t  H (u x  v y )  0 .
(1.3.3)
We compute the vertical vorticity and the horizontal divergence, respectively, from
(1.3.1) and (1.3.2). By utilizing (1.1.3), we then obtain
(v x  u y ) t 
f
t  0 ,
H
(1.3.4)
and
tt
H
 f ( v x  u y )  g ( xx   yy ) .
(1.3.5)
The vorticity equation can be integrated in time, i.e.
vx  u y 
f
f
  v 0 x  u0 y   0 ,
H
H
(1.3.6)
where sub-zeroes denote initial values. We assume that the problem is started from
rest, which means that there are no velocities or velocity gradients at t = 0. Thus
vx  u y  
f
(0   ) .
H
(1.3.7)
Inserting for the vorticity in (1.3.5), we find that
tt  c02 (xx   yy )  f 2  f 20 ,
(1.3.8)
where c02  gH , and 0 is a known function of x and y (the surface elevation at t = 0).
The solution to (1.3.8) can be written as a sum of a transient (free) part and a
stationary (forced) part
  ~( x, y, t )   ( x, y ) ,
(1.3.9)
where ~ and  fulfil, respectively
~tt  c02 (~xx  ~yy )  f 2~  0 ,
(1.3.10)
14
 c02 ( xx   yy )  f 2  f 20 .
(1.3.11)
Equation (1.3.10) for the transient, free solution is called the Klein-Gordon equation
and occurs in many branches in physics. Here, it describes long surface waves that are
modified by the earth’s rotation (Sverdrup waves). These waves will be discussed in
the next section. Notice that the initial conditions for the free solution are
~( x, y,0)   ( x, y )   ( x, y ) and ~  0 .
0
t
As an example of a stationary solution of (1.3.8), we return to the problem in the
last section where the surface elevation initially was a step function:
 h 2 , x  0,
 h 2 , x  0,
0 ( x )  
(1.3.12)
or, for simplicity,
 1, x  0,
 1, x  0.
1
2
0 ( x)  h sgn( x), sgn( x)  
(1.3.13)
We assume that the motion is independent of the y-coordinate. From (1.3.11) we then
obtain
1
2
 xx  a  2   a  2 h sgn( x) ,
(1.3.14)
where a  c0 / f . It is easy to show that the solution of (1.3.14) is given by
1
2
  h{1  exp(  x / a)}sgn( x) .
We have sketched this solution in Fig. 1.5.
Fig. 1.5 Geostrophic adjustment of a free surface.
(1.3.15)
15
The distance a = c0/f = (gH)1/2/f is called the Rossby radius of deformation, or the
Rossby radius for short. A typical value for f at mid latitudes is 104s1. For a deep
ocean (H = 4000 m), we find that a  2000 km, while for a shallow ocean (H = 100
m), a  300 km.
From (1.3.1) and (1.3.2) we find the velocity distribution for this example, i.e.
f v  g x ,
(1.3.16)
u  0.
(1.3.17)
and
We note from (1.3.16) that we have a balance between the Coriolis force and the
pressure-gradient force (geostrophic balance) in the x-direction. From (1.3.15) and
(1.3.16) the corresponding geostrophic velocity in the y-direction can be written
v
gh
exp(  x / a ) .
2c0
(1.3.18)
This is a “jet”-like stationary flow in the positive y-direction. Although the
geostrophic adjustment occurs within the Rossby radius, we notice from (1.3.18) that
the maximum velocity in this case is independent of the earth’s rotation. By
comparison with (1.2.7), we see that our maximum velocity it is the same as the
velocity below a moving pulse with height h/2, or as the velocity in the non-rotating
step-problem.
Let us compute the kinetic and the potential energy within a geometrically fixed
area  D  x  D for the stationary solution (1.3.15)-(1.3.18), which is valid when t
 . The kinetic energy becomes
D
 2 
1
1
Ek      v dz dx   gh 2 a (1  e  2 D / a ) ,


2 D  H
8

(1.3.19)
where we have used the fact that H   . For the potential energy we find
 h / 2 
1
1
E p   g    z ' dz ' dx   gh 2 D   gh 2 a(3  4e  D / a  e  2 D / a ) ,


2
8
D 
0

D
(1.3.20)
where we have taken z = h/2 as the level of zero potential energy, and introduced
z'  z  h / 2 . Initially, the total mechanical energy within the considered area equals
the potential energy, or
E0  E p0 
1
 gh 2 D .
2
Let us choose D >> a. We then notice from (1.3.19)-(1.3.21) that
(1.3.21)
16
Ek  E p  E0 .
(1.3.22)
Thus, when t  , the total mechanical energy inside the considered area is less than
it was at t = 0. The reason is that energy in the form of free Sverdrup waves (solutions
of the Klein-Gordon equation) has “leaked” out of the area during the adjustment
towards a geostrophically balanced steady state. We will consider these waves in
more detail in the next section.
Finally we discuss in a quantitative way when it is possible to neglect the
effect of earth’s rotation on the motion. For this to be possible, we must have that
 

(1.3.23)
vt  fk  v .
Accordingly, the typical timescale T for the motion must satisfy
T 
2
.
f
(1.3.24)
At mid latitudes we typically have 2/f  17 hours. If the characteristic horizontal
scale of the motion is L and the phase speed is c ~ (gH)1/2, we find from (1.3.24) that
the effect of earth’s rotation can be neglected if
L  a ,
(1.3.25)
where a = c/f is the Rossby radius of the problem. In the open ocean L will be
associated with the wavelength, while in a fjord or canal, L will be the width.
Oppositely, when
L  a,
(1.3.26)
the effect of the earth’s rotation on the fluid motion can not be neglected.
1.4 Sverdrup waves, inertial waves and Poincaré waves
We consider long surface waves in a rotating ocean of unlimited horizontal
extent. Such waves are often called Sverdrup waves (Sverdrup, 1927). They are
solutions of the Klein-Gordon equation (1.3.10). Actually, Sverdrup’s name is usually
related to friction-modified, long gravity waves, but here we will use it also for the
frictionless case. In literature long waves in an inviscid ocean are often called
Poincaré waves. However, this term will be reserved for a particular combination of
Sverdrup waves that can occur in canals with parallel walls.
Sverdrup waves
A surface wave component in a horizontally unlimited ocean can be written
17
  A exp( i (kx  ly  t )) .
(1.4.1)
This wave component is a solution of (1.3.10) if
 2  f 2  c02 (k 2  l 2 ) .
(1.4.2)
Here k and l are real wave numbers in the x- and y-direction, respectively, and  is the
wave frequency. Equation (1.4.2) is the dispersion relation for inviscid Sverdrup
waves. From this relation we note that the Sverdrup wave must always have a
frequency that is larger than (or equal to) the inertial frequency f.
For simplicity we let the wave propagate along the x-axis, i.e. l = 0. The phase
speed now becomes

2 

c   c0 1 
2 2 
k
 4 a 

1/ 2
,
(1.4.3)
where  = 2/k is the wavelength and a = c0/f is the Rossby radius. We note that the
waves become dispersive due to the earth’s rotation, i.e. the phase speed depends on
the wavelength (here: increases with increasing wave length). The group velocity
becomes
cg 
c0
d
.

1/ 2
dk 
2 
1 

2 2 
 4 a 
(1.4.4)
We notice that the group velocity decreases with increasing wavelength. From (1.4.3)
and (1.4.4) we realize that ccg = c02, i.e. the product of the phase and group velocities
are constant. From (1.4.2), with l = 0, we can sketch the dispersion diagram for
positive wave numbers; see Fig. 1.6.
Fig. 1.6. The dispersion diagram for Sverdrup waves.
18
For k << a1 (i.e.  >> a) we have that   f. This means that the motion is reduced to
inertial oscillations in the horizontal plane. For k >> a1 gravity dominates, i.e.  
c0k, and we have surface gravity waves that are not influenced by the earth’s rotation.
Contrary to gravity waves in a non-rotating ocean, the Sverdrup waves discussed
here do possess vertical vorticity. For a wave solution (  exp( i t ) ), (1.3.4) yields
 
f
.
H
(1.4.5)
If we still assume that  /y = 0, we obtain from (1.4.5) and (1.3.2) that
f

, 

H
1 
u   vt . 

f
vx 
(1.4.6)
With  in (1.4.1) depending only on x and t, we find, when we let the real part
represent the physical solution:
  0 cos( kx   t ),

u  0 cos( kx   t ),



kH

0 f

v
sin( kx   t ),

kH

zH
w  0 
 sin( kx   t ).
 H 

(1.4.7)
Here the vertical velocity w has been obtained from (1.1.17). Since   f
for
Sverdrup waves, we must have that u  v . Furthermore, from (1.4.7) we find that
u2
v2

 1.
0 /( kH ) 2 0 f /( kH ) 2
(1.4.8)
This means that the horizontal velocity vector describes an ellipsis where the ratio of
the major axis to the minor axis is  / f . From (1.4.7) it is easy to see that the
velocity vector turns clockwise, and that one cycle is completed in time 2/.
Sverdrup (1927) demonstrated that the tidal waves on the Siberian continental
shelf were of the same type as the waves discussed here. In addition, they were
modified by the effect of bottom friction, which leads to a damping of the wave
amplitude as the wave progresses. Furthermore, friction acts to reduce of the phase
speed, and it causes a phase displacement between maximum current and maximum
surface elevation.
19
Inertial waves
When discussing the dispersion relation (1.4.2), we realized that for very long
wavelengths the Sverdrup wave was reduced to an inertial wave. For the inertial wave
phenomenon, the effect of gravity is unimportant. Hence, to study such waves we may
take that the surface is horizontal at all times, i.e.   0. Furthermore, for infinitely
long waves we must put  / x  0,  / y  0 in the governing equations (1.3.1)(1.3.3). They then reduce to
ut  fv  0, 

vt  fu  0. 
(1.4.9)
 d2

 2  f 2 (u, v)  0 ,
 dt

(1.4.10)
Accordingly,
where we have changed to ordinary derivatives since u and v now are only functions
of time. With initial conditions u = u0, v = 0, the solution becomes
u  u0 cos ft, 

v  u0 sin ft. 
(1.4.11)
We notice from (1.4.11) that u 2  v 2  u 02 , i.e. the magnitude of the velocity is u0
everywhere in the ocean, and that the velocity vector turns clockwise in the northern
hemisphere (where f > 0). The trajectory of a single fluid particle (xp, yp) can be found
from the relations u = dxp /dt and v = dyp /dt, where u and v are given by (1.4.11).
Accordingly
u0
sin ft,
f
u
y p  y0  0 cos ft,
f
x p  x0 





(1.4.12)
u02
.
f2
(1.4.13)
or
( x p  x0 )2  ( y p  y0 )2 
This means that a fluid particle moves in a closed circle around the point (x0, y0),
which is the mean position of the particle in question. The radius of the circle is r =
u0/f, which is called the inertial radius. If we use the typical values u0 = 10 cm s1 and
f = 104s1, we find that r = 1 km. The inertial period T = 2/f, which is the time
20
needed for a particle to make one closed loop, is about 17 hours in this example. The
particle moves in a clockwise direction in the northern hemisphere. Since the vertical
component of the earth’s rotation at a location with latitude  is sin, we can
introduce the pendulum day T defined by
T 
2
.
 sin 
(1.4.14)
We notice that the inertial period is half a pendulum day. It turns out that inertial
oscillations are quite common in the ocean. Most ocean current spectra show a local
amplitude or energy maximum at the inertial frequency.
Poincaré waves
We consider waves in a uniform canal along the x-axis with depth H and width B.
Such waves must satisfy the Klein-Gordon equation (1.3.10). But now the ocean is
laterally limited. At the canal walls, the normal velocity must vanish, i.e. v = 0 for y =
0, B. By inspecting (1.4.7), we realize that no single Sverdrup wave can satisfy these
conditions. However, if we superimpose two Sverdrup waves, both propagating at
oblique angles ( and , say) with respect to the x-axis, we can construct a wave
which satisfies the required boundary conditions. The velocity component in the ydirection must then be of the form
v  v0 sin(
n y
) exp( i (kx  t )), n  1,2, 3,..
B
(1.4.15)
Since the wave number l  n / B in the y-direction now is discrete due to the
boundary conditions, the dispersion relation (1.4.2) becomes
1/ 2
 2
n 2 2 
2 2
    f  c0  k  2  , n  1,2,3, ...
B 


(1.4.16)
We notice from (1.4.15) that the spatial variation in the cross-channel direction is
trigonometric. Such trigonometric waves in a rotating channel are called Poincaré
waves. They can propagate in the positive as well as the negative x-direction. We shall
see that this is in contrast to coastal Kelvin waves, which we discuss later in section 1.
In general, the derivation of the complete solution for Poincaré waves is too lengthy
to be discussed in this text. For a detailed derivation; see for example LeBlond and
Mysak (1978), p. 270.
21
Energy considerations
By utilizing the solution (1.4.7), we can compute the mechanical energy
associated with Sverdrup waves. The mean potential energy per unit area of a fluid
column can be written
1
Ep 
2 / 
2 /  
1
 ( g z dz)dt  4 g
0
2
0
.
(1.4.17)
0
The mean kinetic energy per unit area becomes
Ek 
1
2 / 
2 / 

0
1
1 1  f 2 /  2  2
 u 2  v 2  w2 dz}dt  g 
0 ,
2
4 1  f 2 /  2 
H

0
{

(1.4.18)
where we have utilized that kH << 1. We see that in a rotating ocean (f  0), the mean
potential and the mean kinetic energy in the wave motion are no longer equal
(compare with the results (1.2.10) and (1.2.11) for the non-rotating case, where we
have an equal partition between the two). The dominating part of the mean energy is
now kinetic. The total mean mechanical energy becomes
E  Ek  E p 
1
 g02c 2 / c02 .
2
(1.4.19)
Energy flux and group velocity
The formula for the group velocity, c g  d / dk , used in (1.4.4), arises from
purely kinematic considerations. By superimposing two progressive wave trains with
the same amplitude and with wave numbers k and k +k, and frequencies  and 
+, respectively, one finds that the envelope containing the wave crests and wave
troughs, defined as the wave group, will propagate with the speed  / k . However,
the group velocity has also a very important dynamical interpretation. Consider a
Sverdrup wave that propagates along x-axis. This wave induces a net transport of
energy in the x-direction. The energy transport through a vertical cross-section can be
found by computing the mean work done by the fluctuating pressure at that section.
The mean work F , per unit time and unit area, which is performed on the fluid to the
right of a vertical cross-section, can be written
1
F
2 / 
2 / 

0
H
 (  pu dz)dt ,
(1.4.20)
where p is the fluctuating part of the pressure in the fluid. Inserting from (1.1.11) and
(1.4.7), it follows that
22
F
1
 g02c .
2
(1.4.21)
For a regular, infinitely long wave train the mean energy cannot accumulate in space.
Hence there must be a transport of mean total energy E through the considered
cross-section. Denote this transport velocity by ce. Energy balance then requires that
F  ce E .
(1.4.22)
Equation (1.4.22) yields, by inserting from (1.4.19) and (1.4.21), that
ce  c02 / c  c g ,
(1.4.23)
where the last equality follows from (1.4.4). Accordingly, the mean energy in the
wave motion propagates with the group velocity.
The work done by the fluctuating pressure per unit mass and unit time is often
referred to as the energy flux, and the total energy per unit mass as the energy
density. These quantities vary in space and time. Both concepts follow naturally from
the energy equation for the fluid. With no variation in the y-direction, the linearized
equations (1.3.1)-(1.3.3) reduce to
ut  fv   g x , 

vt  fu  0,

t   Hu x . 
(1.4.24)
By multiplying the two first equations by u and v, respectively, and then adding, we
obtain
 1 2


(u  v 2 )   ( gu )  gu x .

t  2
x

(1.4.25)
Obviously, the Coriolis force does not perform any work since it acts perpendicular
to the displacement (or velocity). By inserting from the continuity equation that
u x  t / H into the last term, (1.4.25) becomes
 1  2
g 2  
2
 u  v     ( gu )  0 .

t  2 
H  x
(1.4.26)
We write this equation


ed  e f  0 ,
t
x
(1.4.27)
where the energy density ed and the energy flux ef are defined, respectively, as
1
g
ed  ( u 2  v 2   2 ) ,
2
H
(1.4.28)
23
e f  gu .
(1.4.29)
The mean values for a vertical fluid column become
2 /  
1
1 02 2 
ed 
(
e
dz
)
dt

g
, 
d
2 /  0 H
2 k 2c02 

2 /  
1
1 2
ef 
( e f dz )dt  g0  / k . 

2 /  0 H
2
(1.4.30)
By comparing with (1.4.19) and (1.4.21) we realize that ed  E /  and e f  F /  ,
and hence e f  c g ed . Even though we have only shown this to be valid for Sverdrup
waves, it is a quite general result, and valid for all kinds of wave motion.
1.5 Kelvin waves at a straight coast
We consider an ocean that is limited by a straight coast. The coast is situated at y
= 0; see Fig. 1.7.
Fig. 1.7. Definition sketch.
Furthermore, we assume that the velocity component in the y-direction is zero
everywhere, i.e. v  0. With constant depth and constant surface pressure (1.4.24)
becomes
ut   g x ,
(1.5.1)
fu   g y ,
(1.5.2)
t   Hu x .
(1.5.3)
24
We take that the Coriolis parameter is constant, and eliminate u from the problem.
Equations (1.5.1) and (1.5.2) yield
 yt  f x  0 ,
(1.5.4)
t  ac0 xy  0 .
(1.5.5)
while (1.5.2) and (1.5.3) yield
Here c0  (gH )1 / 2 and a = c0/f. We assume a solution of the form
  G ( y ) F ( x, t ) .
(1.5.6)
Ft
aG'
,

c0 Fx
G
(1.5.7)
By inserting into (1.5.5), we find
where G '  dG / dy . The left-hand side of (1.5.7) is only a function of x and t, and the
right-hand side is only a function of y. Thus, for (1.5.7) to be valid for arbitrary values
of x, y, and t, both sides must equal to the same constant, which we denote by  (  0
for a non-trivial solution). Hence
aG'

G
Ft

c0 Fx
 G  exp( y / a),
 F  F ( x   c0t ).
(1.5.8)
By inserting from (1.5.8) into (1.5.4), we find that
  1 .
(1.5.9)
Accordingly, from (1.5.8), we have solutions of the form
  exp( y / a) F ( x  c0t ) ,
(1.5.10)
  exp(  y / a) F ( x  c0t ) .
(1.5.11)
and
If we have a straight coast at y = 0 and an unlimited ocean for y > 0, as depicted in
Fig. 1.7, the solution (1.5.10) must be discarded. This is because  must be finite
everywhere in the ocean, even when y  . The solution for the surface elevation and
the velocity distribution in this case then become
  exp(  y / a ) F ( x  c0t ),
u
g
exp(  y / a ) F ( x  c0t ).
fa
(1.5.12)
This type of wave is called a single Kelvin wave (double Kelvin waves will be treated
in section 1.8). It is trapped at the coast within a region determined by the Rossby
25
radius. It is therefore also referred to as a coastal Kelvin wave. The Kelvin wave
propagates in the positive x-direction with velocity c0, like a gravity wave without
rotation. The difference from the non-rotating case, however, is that now we do not
have the possibility of a wave in the negative x-direction. This is because the Kelvin
wave solution requires geostrophic balance in the direction normal to the coast; see
(1.5.2). This is impossible for a wave in the negative x-direction in the northern
hemisphere. In general, if we look in the direction of wave propagation (along the
wave number vector), a Kelvin wave in the northern hemisphere always moves with
the coast to the right, while in the southern hemisphere (f < 0), it moves with the coast
to the left; see the sketch in Fig. 1.8 for a single Fourier component in the northern
hemisphere.
Fig. 1.8. Propagation of Kelvin waves along a straight coast when f > 0.
Since the wave amplitude is trapped within a region limited by the Rossby radius, the
wave energy is also trapped in this region. The energy propagation velocity (the group
velocity) is here c g  d / dk  d (c0 k ) / dk  c0 , and the energy is propagating with the
coast to the right in the northern hemisphere. We note that for Kelvin waves the
frequency has not a lower limit (for Sverdrup waves   f).
The oceanic tide may in certain places manifest itself as coastal Kelvin waves of
the type studied here. We will discuss this further in connection with amphidromic
points (points where the tidal height is always zero). From (1.5.12) we notice that the
surface elevation and velocity are in phase, i.e. maximum high tide coincides with
maximum current. It turns out from measurements that the maximum tidal current at a
given location occurs before maximum tidal height. This is due to the effect of
friction at the ocean bottom, which we have neglected so far.
26
The effect of bottom friction will also modify the Kelvin wave solution (1.5.12)
in various other ways. In particular, it turns out that the lines describing a constant
phase (the co-tidal lines) are no longer directed perpendicular to the coast, but are
slanting backwards relative to the direction of wave propagation (Martinsen and
Weber, 1981). This situation is sketched in Fig. 1.9.
Fig. 1.9. Coastal Kelvin waves influenced by friction.
Other frictional effects on Kelvin waves are reduced phase speed (c < c0), and an
amplitude decrease as the wave progresses, which is similar to how friction affects
Sverdrup waves.
1.6 Amphidromic systems
Wave systems, where the lines of constant phase, or the co-tidal lines, form a
star-shaped pattern, are called amphidromies. They are wave interference phenomena,
and in the ocean they usually originate due to interference between Kelvin waves. Let
us study wave motion in an ocean with width B; see the sketch in Fig. 1.10.
Fig. 1.10. Ocean with parallel boundaries (infinitely long canal).
27
Since the ocean now is limited in the y-direction, both Kelvin wave solutions (1.5.10)
and (1.5.11) can be realized. Because we are working with linear theory, the sum of
two solutions is also a solution, i.e.
  exp( y / a) F ( x  c0t )  exp(  y / a) F ( x  c0t ) .
(1.6.1)
In general the F-functions in (1.6.1) can be written as sums of Fourier components. It
suffices here to consider two Fourier components with equal amplitudes:
  Aexp( y / a) sin( kx  t )  exp(  y / a) sin( kx  t ) ,
(1.6.2)
where  = c0k. Along the x-axis, i.e. for y = 0, (1.6.2) reduces to
  2 A sin kx cos  t .
(1.6.3)
This constitutes a standing oscillation with period T = 2/. Zero elevation ( = 0)
occur when
x
n
, n  0,1, 2, ...
k
(1.6.4)
At the locations given by (n/k, 0), the surface elevation is zero at all times. These
nodal points are referred to as amphidromic points.
We consider the shape of the co-phase lines, and choose a particular phase, e.g. a
wave crest (or trough). At a given time the spatial distribution of this phase is given
by t  0 ; i.e. a local extreme for the surface elevation. Partial differentiation (1.6.2)
with respect to time yields that the co-phase lines are given by the equation
exp( y / a) cos( kx   t )  exp(  y / a) cos( kx   t )  0 .
(1.6.5)
We notice right away that the co-phase lines must intersect at the amphidromic points
x  n / k , y  0 for all times. As an example, we consider the amphidromic point at
the origin. In a sufficiently small distance from origin, x and y are so small that we
can make the approximations exp(  y / a )  1  y / a, cos kx  1, sin kx  kx . Equation
(1.6.5) then yields
y  (ka tan  t ) x .
(1.6.6)
This means that the co-phase lines are straight lines in a region sufficiently close to
the amphidromic points. Since tan t is a monotonically increasing function of time in
the interval t  0 to t   /( 2 ) , we see that a co-phase line revolves around the
amphidromic point in a counter-clockwise direction in this example (f > 0). It turns
out that, as a main rule, the co-phase lines of the amphidromic systems in the world
oceans rotate counter-clockwise in the northern hemisphere and clockwise in the
28
southern hemisphere. We notice from (1.6.6) that if we have high tide along a line in
the region x > 0, y > 0 at some time t, we will have high tide along the same line in the
region x < 0, y < 0 at time t   /  , or half a period later.
We now consider the numerical value of  along a co-phase line. Close to an
amphidromic point, (here the origin), we can use (1.6.2) to express the elevation as
y
a
  2 A(kx cos  t  sin  t ) .
(1.6.7)
From (1.6.6) we find that tan t  y /( kax) along a co-phase line. By eliminating the
time dependence between this expression and (1.6.7), we find for the magnitude of the
surface elevation along a co-tidal line:
  2 Ak 2 x 2  y 2 / a 2  .
1/ 2
(1.6.8)
The lines for a given difference between high and low tide are called co-range lines.
These curves are given by (1.6.8), when  is put equal to a constant, i.e.
k 2 x 2  y 2 / a 2  const .
(1.6.9)
We thus see that the co-range lines close to the amphidromic points are ellipses.
In Fig. 1.11 we have depicted co-phase lines (solid curves) and co-range lines
(broken curves) resulting from the superposition of two oppositely travelling Kelvin
waves, both with periods of 12 hours and amplitudes of 0.5 m. The wavelength is 800
km, the width of the channel is 400 km, the depth is 40 m, and the Coriolis parameter
is 104s1. The Rossby radius becomes 198 km in this example. Hence the right-hand
side of the channel is dominated by the upward-propagating Kelvin wave (the one
with minus sign in the phase), and the left-hand side is dominated by the downward
propagating Kelvin wave.
29
Fig. 1.11. Amphidromic system in an infinitely long canal.
1.7 Equatorial Kelvin waves
Close to equator we have that f0  0. The Coriolis parameter in this region can
then be approximated by
f  y ,
(1.7.1)
where the y-axis is directed northwards; see Fig. 1.12.
Fig. 1.12. Sketch of the co-ordinate axes near the equator.
We will find that it is possible to have equatorially trapped gravity waves, analogous
to the trapping at a straight coastline. Assume that the velocity component in the ydirection is zero everywhere, i.e. we assume geostrophic balance in the direction
30
perpendicular to equator. With constant depth, the equations (1.5.1)-(1.5.3) are
unchanged, but now f   y in (1.5.2). By assuming a solution of the form
  G ( y ) F ( x, t ) as before, (1.5.7) becomes
Ft
c G'
 0  .
c0 Fx  yG
(1.7.2)
Accordingly:
F  F ( x   c0t ),
 2
G  exp 
y .
 2c0 





(1.7.3)
By inserting into (1.5.4), we find
  1 .
(1.7.4)
From (1.7.3) we realize that to have finite solution when y   , we must choose 
= 1 in (1.7.4). The solution thus becomes


g
2
2
u  exp(  y / ae ) F ( x  c0t ), 

c0
  exp(  y 2 / ae2 ) F ( x  c0t ),
(1.7.5)
where the equatorial Rossby radius is defined by
ae  (2c0 /  )1 / 2 .
(1.7.6)
We note that the solution (1.7.5), referred to as an equatorial Kelvin wave, is valid at
both sides of equator and that it propagates in the positive x-direction, i.e. eastwards
with phase speed c0 = (gH)1/2. The energy also propagates eastwards with the same
velocity, since we have no dispersion. At the equator  is approximately
21011m1s1. For a deep ocean with H = 4000 m, we find from (1.7.6) that the
equatorial Rossby radius becomes about 4500 km.
Equatorial Kelvin waves are generated by tidal forces, and by wind stress and
pressure distributions associated with storm events with horizontal scales of thousands
of kilometres. When such waves meet the eastern boundaries in the ocean (the west
coast of the continents), part of the energy in the wave motion will split into a
northward propagating coastal Kelvin waves in the northern hemisphere, and a
southward propagating coastal Kelvin wave in the southern hemisphere. Some of the
energy will also be reflected in the form of planetary Rossby waves. This type of
wave will be discussed in section 1.9.
31
1.8 Topographically trapped waves
We have seen that gravity waves can be trapped at the coast or at the equator due
to the effect of the earth’s rotation. Trapping of wave energy in a rotating ocean can
also occur in places where we have changes in the bottom topography. In this case,
however, the wave motion is fundamentally different from that due to Kelvin waves.
While the velocity field induced by Kelvin waves is always zero in a direction
perpendicular to the coast, or equator, it is in fact the displacement of particles
perpendicular to the bottom contours that generates waves in a region with sloping
bottom. We call these waves escarpment waves, and they arise as a consequence of
the conservation of potential vorticity.
Rigid lid
The escarpment waves are essentially vorticity waves. The motion in these waves
is a result of the conservation of potential vorticity. More precisely, the relative
vorticity for a vertical fluid column changes periodically in time when the column is
stretched or squeezed in a motion back and forth across the bottom contours. To study
such waves in their purest form, we will assume that the surface elevation is zero at
all times, i.e. we apply the rigid lid approximation. In this way the effect of gravity is
eliminated from the problem. Let us assume that the bottom topography is as sketched
in Fig. 1.13.
Fig. 1.13. Bottom topography for escarpment waves.
The continuity equation (1.1.20) can now be written as
( Hu ) x  ( Hv ) y  0 .
(1.8.1)
32
Accordingly, we can define a stream function  satisfying
Hu   y , 

Hv   x . 
(1.8.2)
When linearizing, we obtain from the theorem of conservation of potential vorticity
(1.1.25) that
t
 f 
 f 
 u   v    0 .
H
 H x
 H y
(1.8.3)
We here assume that f is constant. Furthermore, we take that H  H ( y ) . By inserting
 from (1.8.2), we can write (1.8.3) as
 2 t 
H
( yt  f x )  0 ,
H
(1.8.4)
where H   dH / dy . We assume a wave solution of the form
  F ( y ) exp( i(kx  t )) .
(1.8.5)
By inserting into (1.8.4), this yields

2
kf H  
 F   k
F  0 .
    
2 
H H  H 
(1.8.6)
This equation has non-constant coefficients and is therefore problematic to solve for a
general form of H(y). We shall not make any attempts to do so here. Instead, we
derive solutions for two extreme types of bottom topography. One of these cases,
where the bottom exhibits a weak exponential change in the y-direction, will be dealt
with in section 10 in connection with topographic Rossby waves. The other extreme
case, where the slope tends towards a step function, will be analysed here; see the
sketch in Fig. 1.14. The escarpment waves relevant for this topography are often
called double Kelvin waves.
Fig. 1.14. The bottom configuration for double Kelvin waves.
33
For trapped waves, the solutions of (1.8.6) in areas (1) and (2) are, respectively
F1  A1 exp( ky), 

F2  A2 exp( ky). 
(1.8.7)
We note that these waves are trapped within a distance of one wavelength on each
side of the step. At the step itself (y = 0), the volume flux in the y-direction must be
continuous, i.e.
v1H1  v2 H 2 ,
y  0.
(1.8.8)
This means that x (and thereby also ) must be continuous for y = 0, i.e. A1 = A2 = A
in (1.8.7). Furthermore, the pressure in the fluid must be continuous for y = 0. The
pressure is obtained from the linearized x-component of (1.1.1), i.e.
px    (ut  fv)  

H
( Hu )t  fHv  .
(1.8.9)
Writing p  P( y ) exp( i (kx  t )) , and applying (1.8.2) and (1.8.5), we find that
 F  f F 
P   

.
H 
 kH
(1.8.10)
By inserting from (1.8.7), with A1 = A2, into (1.8.10), continuity of the pressure at y =
0 yields the dispersion relation
H  H 
1
 f 2
.
H

H
1
 2
(1.8.11)
We note that we always have that   f , and that the wave propagates with shallow
water to the right in the northern hemisphere, i.e.  > 0 when f > 0. These two
properties are generally valid for escarpment waves, even though we have only shown
it for double Kelvin waves with a rigid lid on top.
In the case where the escarpment represents the transition between a continental
shelf of finite width and the deep ocean, this type of waves are often called
continental shelf waves. This kind of bottom topography is found outside the coast of
Western Norway. Here, numerical results show the existence of continental shelf
waves in the area close to the shelf-break, e.g. Martinsen, Gjevik and Røed (1979).
The topographic trapping of long waves near the shelf-break and the currents
associated with these waves, interact with the wind-generated surface waves, which
tend to make the sea state here particularly rough. This is a well-known fact among
fishermen and other sea travellers that frequent this region.
34
The effect of gravity
In general, we must allow the sea surface to move vertically. Let us consider a
wave solution of the form
(u, v, )  exp( i (kx  t )) .
(1.8.12)
For such waves, the linear versions of (1.1.18) and (1.1.19), with constant surface
pressure, yield
g
k  f y ,
f  2
ig
kf   y .
v 2
f  2
u
2





(1.8.13)
We write the surface elevation as
  G ( y ) exp( i (kx  t )) .
(1.8.14)
Inserting into (1.1.20), we find
 2  f 2
fk 
( HG)  
 k 2H 
H  G  0 .

 g

(1.8.15)
For  2  f 2 , i.e. quasi-geostrophic motion, we revert to the gravity-modified
escarpment wave. For f = 0 and H  (tan  ) y , this equation yields the so-called edge
waves. These are long gravity waves trapped at the coast. The trapping in this case is
not due to the earth’s rotation, but is caused by the fact that the local phase speed
gH increases with increasing distance from the coast. If we represent the wave by a
ray, which is directed along the local direction of energy propagation, i.e.
perpendicular to the co-phase lines, the ray will always be gradually refracted towards
the coast. At the coast, the wave is reflected, and the refraction process starts all over
again. The total wave system thus consists of a superposition between an incident and
a reflected wave in an area near the coast. The width of this area depends on the angle
of incidence with the coast for the ray in question. Outside this region, the wave
amplitude decreases exponentially. The edge waves can propagate in the positive as
well as in the negative x-direction. The lowest possible wave frequency is given by
 2  gk tan  .
(1.8.16)
There are also a number of higher, discrete frequencies for this problem; see LeBlond
and Mysak (1978), p. 221. If we take the earth’s rotation into account (f  0), the
35
frequencies for the edge waves in the positive and negative x-directions will be
slightly different.
1.9 Planetary Rossby waves
Like topographic waves, planetary Rossby waves are a result of the conservation
of potential vorticity. However, in this case the relative vorticity of a vertical fluid
column changes periodically in time when the planetary vorticity changes as the
column moves back and forth in the north-south direction. We realize then, that this
motion occurs on a beta-plane. The motion is essentially horizontal. As for
escarpment waves, we can eliminate gravity completely from the problem by
assuming that the surface is horizontal for all times, i.e.   0. The hydrostatic
assumption then yields for the pressure
p   gz  p( x, y, t ) .
(1.9.1)
The linearized equations thus become




1

vt  fu   p y , 


( Hu ) x  ( Hv ) y  0. 


ut  fv  
1
px ,
(1.9.2)
First, we take that H is constant, and that f = f0+y. From the continuity equation, we
can define a stream function  such that
u   y ,
v  x.
(1.9.3)
By eliminating the pressure from (1.9.2) we find
 2H t   x  0 ,
(1.9.4)
where 2H   2 / x 2   2 / y 2 is the horizontal Laplacian operator. This equation also
follows directly from the linearized potential vorticity conservation equation (1.8.3),
when H is constant and fy = . We assume wave solutions of the form
  exp( i (kx  ly  t )) .
(1.9.5)
Inserting (1.9.5) into (1.9.4) yields for the frequency

k
k  l2
2
.
(1.9.6)
36
We shall discuss this dispersion relation in more detail, but first we repeat some
elements of general wave kinematics.

We introduce the wave number vector  defined by




  k1i1  k 2 i2  k 3 i3 ,

and a radius vector r , where




r  r1i1  r2 i2  r3 i3 .
(1.9.7)
(1.9.8)
A plane wave of the type (1.9.5) can now be written
 
 

  A exp( i (  r   t ))  A exp{i   r 

 
t } .
2 

Accordingly, the vectorial phase speed c can be defined by

 
c 2 ,
 2  k12  k 22  k 32 .


The components of the vectorial group velocity c g are given by
cg(1)   / k1 , 

cg( 2 )   / k2 , 
cg( 3)   / k3 . 
(1.9.9)
(1.9.10)
(1.9.11)
In vector notation this becomes
     

c g     ,   i1
 i2
 i3
.
k1
k2
k3
(1.9.12)
If the frequency  only is a function of the magnitude of the wave number vector, i.e.
   ( ) , we refer to the system as isotropic. If we cannot write the dispersion
relation in this way, the system is anisotropic. We now consider the surface in wave
number space given by    (k1 , k2 , k3 )  C , where C is a constant; see Fig. 1.15,
where we display a two-dimensional example.
Fig. 1.15. Constant- frequency surface in wave number space.
37
The gradient  is always perpendicular to the constant frequency surface. From
(1.9.12) we note that this means that the group velocity is always directed along the
surface normal, as depicted in Fig. 1.15. Since the phase velocity is directed along the
wave number vector, e.g. (1.9.10), we realize that if the phase speed and group
velocity should become parallel, then the constant frequency surface must be a sphere
in wave number space. Mathematically, this means that    ( ) , i.e. we have an
isotropic system.
We now return to the discussion of the dispersion relation (1.9.6) for Rossby
waves. Since    (k ,  ) , the system is anisotropic. Hence, the group velocity and the
vectorial phase speed are not parallel. Equations (1.9.6), (1.9.10) and (1.9.11) yield
for the components of the phase speed and group velocity:
c
( x)
c
( y)
k
 k2 
 2  4 ,


l
 kl 
 2  4 ,



(1.9.13)
and
  (k 2  l 2 )

,
k
4
 2 kl

 4 .
l

cg( x ) 
c
( y)
g





(1.9.14)
We note that the x-component of the phase speed is always directed westwards (the xaxis is assumed to be parallel to the equator). However, the energy propagation can
have an eastward or westward component, depending on the ratio l/k. For waves
propagating along latitudinal circles (l = 0), the energy always propagates eastward.
The magnitudes of the phase speed and the group velocity are easily obtained from
(1.9.13) and (1.9.14):
  k 
c  2 ,
  


cg  2 . 


(1.9.15)


We note that c  | c g | . For a typical Rossby wave propagating along equator we take
 = 2/k = 1000 km. With  = 2  10 11 m1s1, the phase speed and period become c 
 0.5 m s1 and T = |2/|  23 days, respectively. We realize that such Rossby
waves are long-periodic phenomena.
38
Since  is negative (we have chosen k > 0), we can define a positive quantity 
by:
 
1
.
2
(1.9.16)
Equation (1.9.6) can then be written as
k   2  l 2
  2.
(1.9.17)
For a prescribed constant frequency , the wave number components k and l are
obtained as points on a circle in the wave number plane. This circle has a radius  and
is centred at (, 0); see Fig. 1.16. The figure also depicts the phase speed and the
group velocity obtained from (1.9.10) and (1.9.12).
Fig. 1.16. Sketch of the relation between the phase speed and the group velocity for
Rossby waves.
We now consider the reflection of Rossby waves from straight coasts in the
north-south direction. Here, we take that the east coast is situated at x = L, where the
x-axis is parallel to the equator; see Fig. 1.17.
Fig. 1.17. Reflection of Rossby waves from a straight coast.
39
As far as wave propagation is concerned, it is the group velocity that represents signal
velocity. Accordingly, for a receiver, or a coast, to experience a physical impulse, it
must have a component of the group velocity directed against it. Accordingly, the
incoming wave has a group velocity with an eastward x-component. This wave is
labelled with subscript 1, i.e.
 (1)  A1 exp( i(k1x  l1 y  1t )) .
(1.9.18)
The reflected wave, labelled with subscript 2, must have a westward-directed group
velocity component. The reflected wave is written as
 ( 2)  A2 exp( i(k2 x  l2 y  2t )) .
(1.9.19)
At the east coast, the normal component of the velocity must vanish, i.e.
 y(1)  y( 2)  0, x  L ,
(1.9.20)
or
il1 A1 exp( ik1L) exp{i(l1 y  1t )}  il2 A2 exp( ik 2 L) exp{i(l2 y  2t )}  0 .
(1.9.21)
If this is to be valid for all y and all t, we must have that
l1  l2 ,


1  2 ,

A1 exp( ik1L)   A2 exp( ik 2 L). 

(1.9.22)
Hence, the frequency and the wave number component in the y-direction are
conserved during the reflection process. This means that the reflected wave number
vector must describe the same circle as the incident one, with the same l; see Fig.
1.18.
Fig. 1.18. Sketch of the relationship between wave characteristics of incident and
reflected Rossby waves.
40
We notice that the group velocity satisfies the law of reflection in ray theory; that is
the angle with respect to the normal of the reflecting plane is the same for the incident
and for the reflected wave. We also note that an incident short wave is reflected from
an east coast as a long wave, while reflection from a west coast (interchange
subscripts 1 and 2 in Fig. 1.18), results in the transformation from a long wave to a
short wave. For this reason the east coast is said to be a source of long waves, while
the west coast is a source of short waves. Since short waves are subject to stronger
dissipation than long waves (dissipation is proportional to the square of the velocity
gradients), this may lead to an increased transition of mean momentum from waves to
currents at the west coasts of the oceans. This has been used (Pedlosky, 1965) as basis
for explaining why the ocean currents are more intense in such regions (western
intensification). We shall return to that question in section 2.
To see how the effect of a free surface modifies the Rossby wave that we have
studied up to now, we consider the linearized version of the vorticity equation
(1.1.22):
 t  f (u x  v y )   v  0 .
(1.9.23)
Eliminating the horizontal divergence u x  v y in this equation by using the linearized
form of (1.1.20), we find
(v x  u y ) t 
f
t   v  0 .
H
(1.9.24)
We now simplify the problem by assuming that the time scale for the Rossby wave is
so large (several days) that the pressure gradient force and the Coriolis force always
have time to adjust to a state of geostrophic balance. This means that the velocity field
in (1.9.24) can be approximated by
v  g x / f ,
u   g y / f ,



(1.9.25)
where f is a mean (constant) value of f. This relation is often called the quasigeostrophic approximation and requires that  2  f 2 . By replacing f with f in
(1.9.24) and inserting from (1.9.25), we obtain
 2Ht    x  a 2t  0 ,
where a  ( gH )1 / 2 / f . By assuming solutions of the form
(1.9.26)
41
  exp( i(kx  ly  t )) ,
(1.9.27)
the dispersion relation is obtained from (1.9.26):
 
k
k  l 2  a 2
2
.
(1.9.28)
If the wavelength is much smaller than the Rossby radius (k 2 , l 2  a 2 ) , the effect of
the surface elevation can be neglected, and we are back to our rigid lid case (1.9.6).
However, if the wavelength is large (k 2 , l 2  a 2 ) , we find from (1.9.10) that
c( x)  
c( y)  
 k 2a 2
k2  l2
 kla 2
k l
2
2

k2
( a ) 2 ,
4

(1.9.29)

 kl
( a ) 2 .
4

(1.9.30)
In this case, the group velocity components become
cg( x )    a 2 ,
c g( y ) 
2  kl
4
( a ) 4 .
(1.9.31)
(1.9.32)
Since a now is a small parameter, we realize that the energy mainly propagates
westward in this case. If the wave number vector is parallel to the equator (l = 0) we
have non-dispersive waves, i.e. c ( x )  cg( x )   a 2 , and c ( y )  cg( y )  0 .
1.10 Topographic Rossby waves
Let us assume that somewhere the relative vorticity is zero. From the theorem of
conservation of potential vorticity (1.1.25) with   0, we find that a displacement
northwards, where f is increasing, generates negative (anti-cyclonic) relative vorticity.
However, we realize that the same effect can be achieved by a northward
displacement if f is constant and the depth H decreases northward. This gives rise to
the so-called topographic Rossby waves. Of course, the existence of such waves does
not require that the bottom does slope in one particular direction.
Topographic Rossby waves are only a special case of escarpment waves when the
bottom has a very weak exponential slope everywhere in the ocean. By letting
H  H 0 exp( y) , equation (1.8.6) reduces to
fk 

F   F    k 2 
F  0 .
 

By assuming
(1.10.1)
42
F  A exp( iy ) ,
(1.10.2)
insertion into (1.10.1) yields the complex dispersion relation
i   k 2   2 
 fk
0.

(1.10.3)
In general we may allow for a very weak change of wave amplitude in the direction
normal to the coast, i.e. we take  in (1.10.2) to be complex:
  l  i .
(1.10.4)
By insertion into (1.10.3), the imaginary part leads to  = /2 (when l  0). From the
real part of (1.10.3) we then obtain

 fk
.
k  l2   2 / 4
2
(1.10.5)
We note that the along-shore phase speed c ( x )   / k is positive. This means that the
wave propagation in the along-shore direction is such that we have shallow water to
the right (in the northern hemisphere).
For a bottom that slopes gently compared to the wavelength (k >> ), we see
from the (1.10.5) that these waves are similar to short planetary waves propagating in
an ocean of constant depth. The expressions for the frequency  are identical, e.g.
(1.9.6), if
   f .
(1.10.6)
This similarity is often used in laboratory experiments in order to simulate planetary
effects. When l = 0, the equations are satisfied for  = 0, i.e. constant amplitude
waves. Such waves propagate along the bottom contours with shallow water to the
right, and mimic planetary Rossby waves along latitudinal circles in an ocean of
constant depth.
1.11 Barotropic instability
Rossby waves in the ocean can be generated by atmospheric forcing, as will be
shown in section 2. They can also develop due to energy transfer from unstable ocean
currents. Assume that we have a stationary mean flow along a latitudinal circle


(1.11.1)
v  U ( y )i ,
and that this initial state is in geostrophic balance, i.e. Py    f U . We disturb
(perturbate) the initial state such that the velocities and the pressure can be written
43
u  U  u~( x, y, t ),
v  v~( x, y, t ),
pP ~
p ( x, y, t ).





(1.11.2)
Here we have assumed that the motion is two-dimensional, i.e. independent of z. If the
perturbations (denoted by ~) are sufficiently small, the equations governing the
motion can be linearized. With w = 0, i.e. constant depth and horizontal surface in the
model depicted in Fig. 1.1, equations (1.1.1) and (1.1.3) reduce to




~
py , 



1
u~t  U u~x  v~U   f v~   ~
px ,

1
v~t  U v~x  f u~  
u~x  v~y  0.

(1.11.3)
By introducing the stream function  for the perturbation, defined by
u~   , v~   , and eliminating the pressure from the equation above, we find
y
x
 2H t  U ( y ) 2H x  (   U ) x  0 .
(1.11.4)
For U  0, (1.11.4) reduces to the equation for Rossby waves (1.9.4). We assume that
the solutions can be written as
   ( y ) exp( ik ( x  ct )) .
(1.11.5)
Furthermore, the ocean is restricted by two straight coasts, defining a channel in the xdirection; see Fig. 1.19.
Fig. 1.19. Horizontal shear flow in an ocean with parallel boundaries.
44
At y = 0 and y = D the normal velocity must be zero, i.e. x = 0. Since this is
valid for all x along the boundaries, we must accordingly have that  = 0 at y = 0, D.
Insertion from (1.11.5) into (1.11.4) yields
 U   

 k 2   0 ,
 U c

   
(1.11.6)
with boundary conditions
  0, y  0, D .
(1.11.7)
If   0, equation (1.11.6) is reduced to the well-known Rayleigh equation for twodimensional perturbations of shear flow in a homogeneous, inviscid fluid.
We assume that the wave number k in (1.11.5) is real, while the phase speed c
and the amplitude function  are generally complex, i.e.
c  cr  ici , 

   r  i i . 
(1.11.8)
From (1.11.5) we see that if ci > 0,  will grow exponentially in time. The initial state
is then said to be unstable, and we have an energy transfer from the mean flow to the
disturbance, which in this case is a Rossby wave.
Since equation (1.11.6) for the perturbation amplitude  has non-constant
coefficients and a singularity for U = c (the term containing the highest derivative
vanishes at locations in the fluid where U = c), it is in general quite complicated to
solve. We will not attempt to do so here. Instead of a detailed investigation of the
stability conditions, we will be content with the derivation of a necessary condition
for the occurrence of instability. First, we make the denominator in (1.11.6) real, i.e.
 U    U  cr  ici 

 k 2   0 .
2
2
U  cr   ci


   
(1.11.9)
Now we multiply this equation with the complex conjugate amplitude function
    r  ii , and note that     r2   i2    0 . By integrating the resulting
2
equation from y = 0 to y = D, and applying the boundary conditions      0 at y =
0, D, we find that
2
2
D
 2

    k 2  2  U    U  cr  dy  ic U     dy  0 .
i
2
0 
2
U  cr 2  ci2 
0 U  cr   ci

D
(1.11.10)
For this equation to be satisfied, both real and imaginary parts must be zero. It
suffices here to consider only the imaginary part, i.e.
45
U     2
dy  0 .
2
2


U

c

c
r
i
0
D
ci 
(1.11.11)
If ci > 0, i.e. the basic flow is unstable, the integral in (1.11.11) must be zero.
However, the integrand is positive, possibly except for the factor U    in the
numerator. Hence, for the integral to become zero, U    must change sign in the
fluid, which means that there must be at least one place where U    .
This
expresses a necessary (but not sufficient) condition for instability of the basic flow.
An equivalent statement is that the absolute vorticity of the basic flow,  abs  f  U  ,
must have an extreme somewhere in the ocean (where d abs / dy  0 ). This kind of
instability is related to the energy transfer from the kinetic energy of the mean flow
(via the velocity shear) to the perturbations, and is called barotropic instability.
Perhaps even more common in the ocean is baroclinic instability, where energy is
transferred from the potential energy of the basic state to the perturbations via the
density stratification. We will return to discuss baroclinic instability in section 3.
Finally, we mention that on an f-plane ( = 0), the necessary condition for
instability is that the basic flow profile must have a point of inflexion ( U   0 )
somewhere in the fluid. Also, the numerical value of the vorticity U  must have a
maximum where U   0 (the Rayleigh-Fjørtoft criterion). The last condition can be
found by applying that the real part of (1.11.10) is zero, together with (1.11.11) and ci
 0.
1.12 Barotropic flow over an ocean ridge
We consider an ocean current that crosses a sub-sea ridge. For simplicity, we
assume that the ridge is infinitely long in the y-direction, and that the depth far away
from the ridge is constant (= H0); see Fig. 1.20,
46
Fig. 1.20. Model sketch.
The surface far upstream is horizontal and the constant upstream flow U0 is
geostrophic, i.e.
fU 0  
1 Ps
.
 y
(1.12.1)
On an f-plane, which we assume here, this means that the surface pressure Ps of the
basic flow must vary linearly with y.
We study the effect of the ridge by applying the potential vorticity theorem
(1.1.25). In a steady state we have

v  Q  0 .
(1.12.2)
A possible solution of this equation is Q  const. , in the (x, y)-plane. We assume that
the effect of the ridge is not felt far upstream, i.e. Q  f / H 0 for all values of y when
x   . Since Q is a conserved quantity, we then must have
Q
f 
f

H  H0
(1.12.3)
We introduce the shape of the ridge q(x), i.e.
H  H 0  q ( x) .
(1.12.4)
The typical width of the ridge is 2L; see Fig. 1.20. Inserting (1.12.4) into (1.12.3), and
assuming that  / y  0 , we obtain
vx 
f
(  q) .
H0
(1.12.5)
47
In the literature one often finds discussions of this problem where the surface is
modeled as a rigid lid, i.e.   0 . In this case we notice from (1.12.5) that anticyclonic vorticity is generated over the entire ridge. By integration we obtain
x
v
f
q ( )d .
H 0 L
(1.12.6)
For x  L , we find that v = 0. For x  L , we obtain
v  v ( )  
fA
,
H0
(1.12.7)
where A is the cross-section of the ridge, given by
A

L

L
 q( )d   q( )d .
(1.12.8)
The corresponding streamlines are sketched in Fig. 1.21.
Fig. 1.21. Deflection of flow over a ridge when the surface is approximated by a rigid
lid.
The deflection angle  of the streamlines on the lee side of the ridge is
 v ( ) 
 fA 
  arctan 
 .
 U0 
 U0H0 
  arctan 
(1.12.9)
The solution (1.12.6) may not be realistic, since it (through geostrophic balance in the
x-direction) implies that p / x must be constant far downstream. This means that the
perturbation pressure becomes infinitely large when x   . However, the solution
obtained here may be of interest close to the ridge.
In most geophysical applications the surface can move freely in the vertical
direction. This means that a fluid column can be stretched, and not just be equal to, or
48
shorter than the upstream value, as in our previous example. From the linearized
versions of (1.1.18)-(1.1.19), we obtain for the surface elevation that is induced by the
ridge:
x 
U 02
f
v xx  v .
gf
g
(1.12.10)
By insertion into (1.12.5), we find
(1  F02 )v xx 
1
f
v
qx .
2
a
H0
(1.12.11)
Here F0  U 0 /( gH 0 )1 / 2 is the Froude number, and a  ( gH 0 )1 / 2 / f is the Rossby
radius. In the ocean F0 << 1, while we typically have that a  103 km. It is then
reasonable to assume that L << a. Hence, on the left-hand side of (1.12.11), the first
term dominates. Accordingly, a particular solution, when F0 << 1, can be written
v( p)  
f
qdx  C1 x  C 2 .
H0 
(1.12.12)
The solution of the homogeneous equation is:
v ( h )  C3 exp( x / a)  C4 exp(  x / a ) .
(1.12.13)
The coefficients C1, C2, C3, and C4 are determined from the boundary conditions. We
assume that the ridge is symmetric about x = 0, and we require that the velocity and
the pressure are continuous at x = 0. Furthermore, we assume that v  0, x   ,
and that the perturbation pressure becomes finite when x   . The last condition is
equivalent to require vanishing v at infinity (in the case of no waves behind the ridge).
The solution becomes:
x
 fA
f
exp( x / a) 
q( )d , x  0,

H 0 
 2H 0
v
x
 fA {1  exp(  x / a)}  f q( )d , x  0.
 2H
H 0 0
 0
(1.12.14)
Here A is the cross-sectional area defined by (1.12.8). This solution yields v (0)  0 .
For x < L, we find that v  fAexp( x / a ) /( 2H 0 )  0 , while for x > L we have that
v   fAexp(  x / a ) /( 2 H 0 )  0 . For a symmetrical ridge, v becomes anti-symmetric
about the centerline of the ridge. The streamlines in this case are sketched in Fig.
1.22.
49
Fig. 1.22. Deflection of flow over a symmetric ridge when the surface is free to move.
In this case we only have a local disturbance of the current in the area close to the
ridge. This is also obvious from the conservation of potential vorticity (1.12.3):
f 
f

.
H0  q  H0
(1.12.15)
We have stretching of a fluid column, and an associated generation of cyclonic
vorticity just before, and just after the ridge. Over the higher parts of the ridge the
column becomes compressed, and here anti-cyclonic vorticity is generated; see the
schematic sketch in Fig. 1.23.
Fig. 1.23. Conservation of potential vorticity for a fluid column when the upstream
fluid depth H0 is independent of the y-coordinate.
In the present case we will always have an elevation of the surface over the ridge, as
sketched in Fig. 1.23. By inserting from (1.12.5) into (1.12.14), we find that
 ( 0) 
A
0.
2a
(1.12.16)
50
The elevation computed here is caused by the effect of the earth’s rotation. In the
general case where the Froude number is not small, we find from (1.12.5) and
(1.12.10) that
 xx 
F02
1
1



q  2
q.
2
2
2 xx
a (1  F0 )
1  F0
a (1  F02 )
(1.12.17)
Solutions of this equation may yield surface elevations or surface depressions above
the ridge, depending of the form of q, and the magnitudes of a and F0. We will not
discuss the general solution here. However, for a non-rotating fluid (f = 0, i.e. a),
we find from (1.12.17) that

F02
q( x ) .
1  F02
(1.12.18)
This is a well-known result in fluid mechanics. It states that for sub-critical currents
(F0 < 1), we have  (0)  0 above a ridge. For super-critical currents (F0 > 1), we find
 (0)  0 . Examples of both these cases are found for flow over bottom rocks in
shallow rivers.
Stationary Rossby waves behind ridges
For large-scale motion over wide ridges we must take into account the fact that
the Coriolis parameter varies with the latitude, i.e. f  f 0   y . Then stationary, or
standing, planetary Rossby waves can be generated by the presence of the ridge. In
general, the condition for the existence of a standing wave is that it must have a phase
speed that is equal in magnitude, and oppositely directed to the basic current. The
direction of the energy flow (basic current plus group velocity), then decides on which
side of the ridge we find the standing wave, because non-damped wave motion cannot
exist without continuous supply of energy to the wave region. In our case, the phase
of the Rossby wave moves westward, while the energy moves eastward. This apply
to planetary Rossby waves where the wavelength  and the Rossby radius a satisfy
the relation   2 a ; see the dispersion relation (1.9.28). We then realize
immediately that standing Rossby waves may occur on the lee-side of the ridge (here:
for x > 0), if the current is directed from the west towards the east. Furthermore, it is
obvious that we cannot have standing Rossby waves at all when the current is directed
towards the west. In the atmosphere, standing, mountain-generated Rossby waves in
51
the westerlies may be responsible for observed distributions of pressure ridges and
pressure troughs; see Charney and Eliassen (1949).
Alternatively, on an f-plane, we may find standing topographic Rossby waves on
the lee-side of the ridge in our example (Fig. 1.20) if the ocean bottom slopes
upwards in the y-direction (remember: decreasing H gives the same effect as
increasing f in the potential vorticity, e.g. sec. 1.10). If H  H 0 exp( y) , where  < 0,
the frequency for short Rossby waves is    f / k  0 , e.g. (1.10.5). As mentioned
before, a standing wave requires a phase speed that is equal, and oppositely directed
to the basic flow, i.e.

k
 U0 .
(1.12.19)
Accordingly, the wavelength of standing topographic Rossby waves becomes
U
  2  0
 f
1/ 2




.
(1.12.20)
This scale is relevant for laboratory simulations of Rossby waves in rotating tank
experiments.
When ocean currents are flowing over isolated sub-sea mountains or banks, anticyclonic vorticity may be generated over the top of the mount. If the vorticity is
sufficiently strong, as it can be for high mounts, a permanent vortex, or a Taylor
column (Taylor, 1922) may be formed. Measurements from the Halten Bank west of
Norway (Eide, 1979), suggest the presence of such vortexes, or eddies. This means
that the water mass above the bank may have a rather long residential time. But bank
regions are important for the early growth of fish larvae. Therefore, in the case of
accidental oil spills, or other kind of pollution, the long residence time of the water
mass means that the fish larvae will stay in contact with the polluted water for quite
some time.
2. WIND-DRIVEN CURRENTS AND OCEAN CIRCULATION
2.1 Equations for the mean motion
Mass and momentum
The velocity introduced in Section 1 was related to the displacement of individual
52
fluid particles. However, many oceanic flow phenomena are turbulent, which means
that it is difficult to keep track of individual particles. But turbulent flows are not
without structure. Even if individual particles move in a chaotic way, they may in an
average sense define an orderly flow. In the ideal case the mean flow has a much
larger characteristic time scale than the rapid fluctuations of individual fluid particles.
Then we can define mean velocities by a time averaging process:

1
v
T
t  T / 2

 v dt .
(2.1.1)
t  T / 2
Here T represents a period that is large compared to the characteristic scale of the
turbulence, but small compared to the typical time scale of the mean motion we intend
to study. The same type of averaging must be done for the pressure and the density.
Accordingly we write
  
v  v  v , 

     , 
p  p  p. 
(2.1.2)

where v  ,   , p  represent the fluctuating (turbulent) part of the motion.
The continuity equation (1.1.2) expresses the conservation of mass in a fluid, and
is a fundamental law in non-relativistic mechanics. Obviously, molecular processes
like heat and salt diffusion may change the density of a fluid particle. This effect is
accounted for by the application of the equation of state for the fluid. The change of
density has a very important dynamical implication in that it changes the buoyancy of
a fluid particle, which in turn may lead to thermohaline circulations in the system.
Equation (1.1.2) is derived under the assumption that only advective fluxes (transport
of mass by the velocity field) can change the mass within a fixed, small geometrical
volume. This assumption is essential for the definition of the velocity field in the
fluid.
It is a fact that for most low-frequency phenomena in the ocean, the individual
change of density for a fluid particle is very small, and can be neglected in the
continuity equation. Hence, we take that

v  0,
(2.1.3)
i.e. the velocity field in the ocean is approximately non-divergent. By averaging this
equation and applying (2.1.2), we find that
53

  v  0, 


  v   0. 
(2.1.4)
By separating the variables as in (2.1.2), and applying (2.1.4) we find from averaging
the viscous Navier-Stokes equation in a reference system rotating with constant
angular velocity f/2 about the vertical axis, that


 

  
 
1
vt  v  v  fk  v    p  gk    v  v v  .

(2.1.5)
We have neglected   compared to  in the pressure term, since typically in the ocean
  /  ~ 103 . Furthermore,  is the kinematic molecular viscosity coefficient (

102 cm2s1). The term v is usually much smaller than the turbulent momentum


transport per unit mass,  v v  , and can therefore be neglected. The term   v v  is
usually referred to as the turbulent Reynolds stress tensor.
We first consider the case of isotropic turbulence, i.e. turbulence that has no
preferred directions, and introduce a constant kinematic eddy viscosity A0. Analogous
to the stress tensor for a Newtonian fluid, we assume for the turbulent Reynolds stress
tensor per unit density:
 v v j  2
 viv j  A0  i 
   ij K , i, j  1,2,3 .
 x j xi  3
(2.1.6)
Here
1, i  j ,
0, i  j ,
 ij  
and
K
1 3
 vivi .
2 i 1
The quantity K is the kinematic energy per unit mass of the turbulent fluctuations. The
sum of the diagonal, turbulent Reynolds stress components (i = j) must always equal
2K. We note from the right-hand side of (2.1.6) that this is true for our choice of

parameterisation, since   v  0 . The term 2 K / 3 formally acts as an “isotropic”
pressure. This term can be combined with the mean pressure. In the momentum
equation we then write
1
2
1
2
1
 p  K  ( p  K )   p* ,

3

3

(2.1.7)
neglecting the variation of  in connection with K. If we assume that A0 is much
larger than the molecular viscosity coefficient, insertion from (2.1.6) and (2.1.7) into
(1.2.5) yields
54
 

  

1
vt  v  v  fk  v    p*  A0 2v  gk .

(2.1.8)
Here 2   2 / x 2   2 / y 2   2 / z 2 is the three-dimensional Laplacian operator.
For large-scale ocean circulation it is usual to assume that the turbulence is
dependent on direction, and we define AH and A as turbulent eddy viscosity
coefficients in the horizontal and the vertical directions, respectively. The surfaces of
constant density are mainly horizontal in the ocean, and the stratification is stable.
Hence the turbulence activity in the vertical direction will be smaller than in the
horizontal direction. Therefore one usually assumes that AH >> A. For such motion wz
is often negligible, i.e. u x  v y  0 . Furthermore, we typically have u2  v 2  w2 .
Accordingly, we can assume for the normal stresses:
 uu  2 AH u x  K , 

 vv  2 AH v y  K , 

 ww  0,

(2.1.9)
while the tangential stresses become:
 uv  vu  AH (vx  u y ), 

 uw   wu  Au z  AH wx , 
 vw   wv  Avz  AH wy . 
(2.1.10)
Furthermore, we assume that AH is constant, while we permit that A  A(z ) . The
horizontal component of (2.1.5) can then be written
 
  


1
qt  v  q  fk  q    H p*  AH  2H q  ( Aqz ) z ,

(2.1.11)
 

where q  ui  vj ,  2H   2 / x 2   2 / y 2 , and p*  p  K .
For large-scale ocean circulation we can assume that the vertical pressure
distribution is approximately hydrostatic. This means that we can replace the zcomponent in (2.1.5) by

p  g   dz  PS ( x, y, t ) ,
(2.1.12)
z
where PS is the atmospheric pressure at the sea surface.
Heat and salt
The internal energy per unit mass of sea water, e, can be written approximately as
55
e  cv T ,
(2.1.13)
where cv is the specific heat at constant volume, and T the absolute temperature. By
considering a geometrically fixed volume  with surface , we find for the change of
internal energy with time within  that





(2.1.14)

e



(

e
v
)




F
T   ,



t 



where the vectorial area  is normal to the surface and points outwards from the
considered fluid volume. The two terms on the right-hand side represent the advective
and diffusive surface heat fluxes, respectively, into the volume. We assume that the

diffusive energy flux, FT , is given by Fourier's law:

FT  kT T .
(2.1.15)
Here kT is the coefficient of heat conduction. Applying Gauss’ theorem, we obtain
from (2.1.14)



  t (  c T )    (  c T v )    (k T )   0 .
v
v
T
(2.1.16)
This equation is valid for an arbitrary volume  in the fluid. Accordingly, the
integrand must be zero everywhere. Furthermore, by taking that cv and kT are
constants, which are reasonable assumptions, and using (2.1.3), we find from (2.1.16):
T 
 v  T   T  2 T .
t
(2.1.17)
Here  T  kT /(  cv ) is called the thermal diffusivity. Equation (2.1.17) is often
referred to as the heat equation.
Equivalently, the change of salt with time within a geometrically fixed volume 
is given by






S



(

S
v
)




F
S   ,



t 


(2.1.18)
where the terms on the right-hand side represent the advective and diffusitive surface
salt fluxes, respectively, into the fluid volume. We assume that the diffusitive salt

flux, FS , is given by Fick’s law

(2.1.19)
FS  k S S ,
where kS is the diffusion coefficient for salt in water. As for heat, we now obtain for
salt from (2.1.18)
56



  t ( S )    ( Sv )    (k S )  0 .
(2.1.20)
S
If we assume that  and kS are approximately constants in this equation, and apply
(2.1.3), we find that
S 
 v  S   S  2 S .
t
(2.1.21)
Here  S  k S /  is called the salt diffusivity. Equation (2.1.21) is often referred to as
the salt equation.
The diffusion coefficients T and S are determined by the molecular structure of
the fluid. They vary in fact with the temperature and the salinity, but this variation is
small, and we will regard them as constants. Typical values for seawater are T 
103cm2s1 and S  105cm2s1.
We define T  T  T  and S  S  S  . Furthermore, we introduce turbulent eddy
diffusivities K(T) and K(S) for temperature and salt. The eddy diffusivities are generally
much larger than the molecular diffusivities in (2.1.17) and (2.1.21). Analogous to the
molecular fluxes from Fourier’s and Fick’s laws, we now assume for the turbulent
transports
 uT   K H(T )Tx ,  vT   K H(T )Ty ,  wT   K (T )Tz ,
 uS   K H( S ) S x ,  vS   K H( S ) S y ,  wS   K ( S ) S z .



(2.1.22)
As before, subscript H denotes horizontal direction. We assume that the horizontal
eddy diffusivities are constant. Hence the heat and salt equations become, when we
neglect the effect of molecular diffusion:

Tt  v  T  K H(T )  2H T  ( K (T )Tz ) z ,
(2.1.23)

S t  v  S  K H( S )  2H S  ( K ( S ) S z ) z .
(2.1.24)
In fully developed turbulent motion the exchange of heat and the exchange of salt
must be approximately similar, i.e.
K H(T )  K H( S ) , 

K (T )  K ( S ) . 
(2.1.25)
However, the mechanisms for momentum exchange in turbulent flows may be
different from those related to heat/salt, so that we generally have AH  KH and A  K.
In the upper mixed layer in oceans and fjords the motion sometimes seems to be
quasi-turbulent. In such cases we find that K(T) > K(S) (Aas, private communication).
57
To get a closed system, the equation of state must be specified. The general form
is
  F (T , S , p) .
(2.1.26)
For seawater, F is a very complicated function. Usually we can replace the variables
in F with its mean values, i.e.
  F T , S , p  .
(2.1.27)
We have now seven equations, (2.1.4), (2.1.11), (2.1.12), (2.1.23), (2.1.24), (2.1.27)
for the seven unknowns, u , v , w , p,  , T , S , which vary in space and time. When we in
the following use some of these equations to describe the mean motion, the mean sign
will be left out for simplicity. We also replace p* by p , i.e. we neglect the
contribution from the turbulent Reynolds stresses to the mean pressure.
Along the sea surface we have wind stress components  S( x ) , S( y ) and a variable air
pressure PS ( x, y, t ) that act on the ocean. The wind stress components are related to
the wind speed (U10, V10) at 10 meters through the semi-empirical relations
 S( x )  a cD (U102  V102 )1 / 2U10 ,

 S( y )  a cD (U102  V102 )1 / 2V10. 
(2.1.28)
where a and cD are the density of air (1.2103g cm3) and the drag coefficient,
respectively. Attempts have been made to assess the value of cD from measurements,
and the results vary considerably. Values between 1103 and 3103 are reported in
the literature, where higher values correspond to increasing wind speeds.
2.2 The Ekman elementary current system
We consider stationary motion in a quasi-homogeneous ocean. By quasihomogeneous we mean that the generally varying density in the ocean can be
approximated by a constant value in these studies. The ocean extends to infinity in the
horizontal directions, and has constant depth H. We assume that the ratio between the
convective acceleration and the Coriolis force, defining the Rossby number, R, is
small, i.e.
 
 
R  v  v H / fk  v H  1 .
(2.2.1)
This allows us to disregard the acceleration term in the momentum equation. We also
assume that horizontal turbulent diffusion of momentum can be neglected. Equation
(2.1.11) for the mean variables (skipping the over-bar) then reduces to
58

px  ( Auz ) z , 



1
fu   p y  ( Avz ) z . 


 fv  
1
(2.2.2)
By assuming constant pressure gradients and constant A in the entire fluid, (2.2.2) has
an exact solution (e.g. Krauss, 1973, p. 248). However, friction is only important in
relatively thin layers close to the surface and the bottom. These are the so-called
Ekman layers, with thickness S and B, respectively. In the following we let the
indices S and B signify the upper and lower Ekman layer, respectively. In these layers
we assume that the eddy viscosities AS and AB are constant, but differing in numerical
value. Furthermore, we assume that H >> S, B, implying that we have a large
interior where the effect of friction can be completely neglected. Here we have
geostrophic balance. Assuming that the horizontal pressure gradient is known, i.e.

 H p  g H   H PS from (2.1.12), the geostrophic velocity v G is given by

px , 



1
fuG   p y . 


 fvG  
1
(2.2.3)
We assume here that f is constant. Then we note from (2.2.3) that the geostrophic flow
is non-divergent:
u xG  v Gy  0 .
(2.2.4)
The total steady flow is subdivided into an Ekman part and a geostrophic part:
 

(2.2.5)
v  v E ( x, y, z )  v G ( x, y) .
At the surface the tangential stress in the fluid must equal the wind stress given by
(2.1.28). This leads to
u
  S( x ) ,
z
v
 AS
  S( y ) .
z
 AS


 z  0.


(2.2.6)
In general, we will allow for a slipping motion at the bottom. For simplicity we
assume that the frictional stress at the bottom is proportional to the bottom velocity
(Rayleigh friction). Hence
59
u B

  r uB , 
z
z  H ,
(2.2.7)

vB
( y)

 B   AB
  r vB ,
z




where r is a friction coefficient and v B  v BE  v BG . Formally, (2.2.7) yields the

extreme conditions “no-slip”, or vB  0 in the limit r   , and “free-slip”, or

vB / z  0 when r = 0.
 B( x )   AB
We start out by studying the Ekman layer at the bottom, and we define a
boundary-layer coordinate ~z such that
~
z  ( H  z) /  B ,
(2.2.8)
where  B   (2 AB / f )1 / 2 is the thickness of the Ekman layer at the bottom. By
introducing the complex velocity W  uBE  iv BE , the equation for the Ekman current
in the bottom layer becomes
W~~  2 2 iW  0 .
zz
(2.2.9)
The boundary conditions at bottom can be written

 AB  (uBE  u G )
E
G



u

u
,

B
~
z

 r B 

E
G
 AB   (vB  v )
E
G 


 vB  v .

~
z
 r B 
~
z  0,
(2.2.10)


Furthermore, above the bottom layer, we must have that v B  v G . Accordingly, in
our new coordinate:

v BE  0,
~
z  .
(2.2.11)
A solution of (2.2.9) that satisfies (2.2.11), can be written
W  C exp{ (1  i )~
z} ,
(2.2.12)
where C = C r+ iCi is complex. By insertion into (2.2.10), we find

1
( K  1)u G  KvG
2K  2K  1
1
Ci 
Ku G  ( K  1)v G ,
2
2K  2K  1
Cr  
2


, 



(2.2.13)
where K  ( AB f / 2)1 / 2 / r . Now, the total velocity in the bottom Ekman layer can be
written
uB  u G  Cr cos  ~
z  Ci sin  ~
z exp(  ~
z ), 

G
~
~
~
vB  v  Ci cos  z  Cr sin  z exp(  z ). 
(2.2.14)
60
For simplicity, we take that u G  0 , and let r (or K0), i.e. we apply a no-slip
condition at the bottom. We then find

u B  v G sin  ~
z exp(  ~
z ),

vB  v G (1  cos  ~
z exp(  ~
z )). 
(2.2.15)
From this result we see that the flow in the bottom Ekman layer (the Ekman spiral) is
veering to left of the direction of the geostrophic current as we move towards the
bottom. For the other extreme case at the bottom, when K (free slip), we find that


Cr = Ci = 0, and accordingly v B  v G . We realize that we need a frictional force at the
bottom to obtain an Ekman spiral. By choosing AB = 10 cm2s1 and AB = 100 cm2s1 as
examples, we find that B becomes 14 m and 44 m, respectively. With r = 0.24 cm s1,
which often is used for shallow oceans, we find that K = 0.1 and 0.3. It could be
realistic to expect that K  0.3 in the ocean.
Taylor (1916) solved the analogous problem for the atmosphere. He assumed that
the frictional stress at the ground was proportional to the square of the wind speed,
instead of using the simplified model with linear friction, as assumed here.
In the upper Ekman layer we define a boundary-layer coordinate
z  z /  S ,
(2.2.16)
where  S   (2 AS / f )1 / 2 . By introducing the complex velocity W  uSE  iv SE for the
Ekman flow, we now find that
Wz z   2 2iW  0 .
(2.2.17)
The boundary conditions (2.2.6) can now be written
uSE
 T ( x) ,
z
vSE
 T ( y) ,
z





z  0,
(2.2.18)
where
T
( x)
 2
 
  AS f
1/ 2

 2
  S( x ) , T ( y )  
  AS f

1/ 2

  S( y ) .

(2.2.19)
We have assumed that H >> S. Accordingly, for the Ekman current
u SE , v SE  0,
z    .
(2.2.20)
The equations (2.2.17)-(2.2.20) define the traditional Ekman spiral for a deep ocean.
The solution is
61


1

(T ( x )  T ( y ) ) cos  z  (T ( x )  T ( y ) ) sin  z exp( z), 
2

1
E
vS  
(T ( x )  T ( y ) ) cos  z  (T ( x )  T ( y ) ) sin  z exp( z). 
2

uSE 


(2.2.21)
This solution describes the oceanic surface Ekman spiral veering to the right (in the
northern hemisphere) compared to the wind direction. The surface current is deflected
45 to the right of the wind. The total current in the surface layer becomes
uS  uSE  u G , 

vS  vSE  vG . 
(2.2.22)
The oceanic flow described in this section is driven by wind stresses and pressure
gradients, and is referred to as Ekman’s elementary current system. The pressure
gradients induce a geostrophic flow, which is felt from the top to the bottom
(independent of z). Close to the ocean bottom this flow is modified by friction in the
Ekman layer. As result, the flow here is veering to the left of the geostrophic flow
direction. The wind stress is only felt in the surface Ekman layer, inducing a current
that is veering to the right of the wind direction.
2.3 The Ekman transport
The volume transport in the lower Ekman layer is defined by

~

u
dz


u
d
z
,
B
B
B
 H
0


 H  B
1

VB   vB dz   B  vB d~
z, 
H
0

UB 
 H  B
1
(2.3.1)
where ~z is defined by (2.2.8). As before, we take u G  0 . By inserting from (2.2.14)
into (2.3.1), we find
2K  1
 BvG ,
2
2 (2 K  2 K  1)


1
G
VB  1 
  Bv ,
2
2

(
2
K

2
K

1
)


UB  





(2.3.2)
where we have utilized that exp(  )  1 . The volume transport is deflected to the
left of the geostrophic flow in the interior. We define a deflection angle B as
 B  arctan
UB
.
VB
(2.3.3)
For K = 0 (no-slip), B = 11.3, while for K = 0.3 (corresponding to AB = 100 cm2s1),
62
we find B = 9.4. Therefore, it is reasonable to assume that the volume transport in
the bottom Ekman layer is deflected about 10 to the left of the geostrophic flow. We
note that this result is derived by using a constant AB in the Ekman layer.
In the upper layer, the Ekman transport is independent of the variation of the
vertical eddy viscosity. Applying constant pressure in (2.2.2) and integrating over the
boundary layer, while using the boundary conditions (2.2.6), we find that

dz   S( y ) /(  f ), 

 S

0
E
E
( x)
VS   vS dz   S /(  f ). 

 S

0
U 
E
S
 u
E
S
(2.3.4)
We notice the well-known result that the Ekman transport in the surface layer is
deflected 90 to the right of the wind direction (in the northern hemisphere). The
result (2.3.4) is based on the fact that the viscous stresses in the fluid are
approximately zero for z  S, which can be seen to be true from (2.2.21). However,
we must remember that (2.2.21) was derived by assuming a constant eddy viscosity in
the upper layer, while this is not done for the Ekman transport. When the eddy
viscosity varies in the upper layer, a precise definition of S in (2.3.4) is not entirely
obvious. However, we may take that  S   (2 AS / f )1 / 2 , where AS is some kind of
average (bulk) value for the upper layer.
2.4 Divergent Ekman transport and forced vertical motion (Ekman suction)
We assume that the geostrophic flow and the wind stress vary slowly in space
and time. By integrating the continuity equation (2.1.4) for the mean flow vertically
across the bottom Ekman layer, we find
wB  
U B VB
,

x
y
(2.4.1)
where wB is the vertical velocity at z   H   B . Furthermore, we have utilized that
w(H) = 0 (horizontal bottom). By inserting (2.2.14) into (2.3.1), we can calculate UB
and VB. To simplify, we assume that K = 0. Applying (2.2.4), we then finally obtain
from (2.4.1)
wB 

B 

k  (  v G )  B  G ,
2
2
(2.4.2)
where  G is the geostrophic vorticity, and we have neglected exp() in comparison
63
with 1.
A similar integration of the continuity equation in the upper Ekman layer yields
wS 
U S VS

1 


k  (   S ) ,
x
y
f
(2.4.3)
where the final result follows from (2.3.4). Here wS is the vertical velocity at z   B .
Furthermore, we have neglected any vertical motion of the sea surface, i.e. we have
taken w(0) = 0. The situation is sketched in Fig. 2.1.
Fig. 2.1. Schematic model of Ekman suction.
As an example, we let a wind field with maximum wind speed of 40 knots ( 20 m
s1) vary over a typical distance L = 100 km. From (2.1.28) we then estimate that S
/(L) ~ 106cm s2. With f = 104s1, equation (2.4.3) yields wS ~ 0.01cm s1. This, in
turn, means that this wind field must be “turned on” for 11½ days to transport a
particle 100 m vertically. In comparison, we can estimate the turbulent vertical
diffusion velocity wturb. For a typical vertical length scale H and a time scale T, the
diffusive terms balance; see for example (2.1.11), when H/T ~ A/H. This defines a
turbulent diffusion velocity scale, i.e. wturb ~ A/H. With A = 100 cm2s1 and H = 100
m, we find that wturb ~ 102 cm s1, which is the same order of magnitude as for the
forced vertical suction velocity in this example.
Let us now assume that the wind stress is zero. Then wS = 0 from (2.4.3). In the
interior of the fluid, where we can disregard the effect of friction, the linearized
vorticity equation from (1.1.18)-(1.1.19) for non-divergent flow on an f-plane reduces
to
 t  f wz .
Vertical integration in the interior yields
(2.4.4)
64
 S
  dz   f w
t
B
,
(2.4.5)
 H  B
since we have assumed that wS = 0. Because of the vertical transport of fluid particles
in the interior, the geostrophic vorticity in this region changes slowly in time, i.e.
 G  F (t )  ( x, y) .
(2.4.6)
We assume that  t   tG  Ft  in equation (2.4.5). By utilizing that H >> S, B, we
approximately have from (2.4.5):
H tG   f wB  
f B G
 ,
2
(2.4.7)
where the last result follows from (2.4.2). Inserting from (2.4.6):
Ft  
f B
F.
2H
(2.4.8)
This equation has a simple exponential solution. The slowly-changing geostrophic
vorticity can then be written
 G   ( x, y ) exp( 
f B
t) .
2H
(2.4.9)
We notice that the vorticity decreases in time. This phenomenon is referred to as spindown, and is due to the vertical transport of fluid particles caused by divergence in the
bottom Ekman layer. In this way the friction, which only is important in a relatively
thin region close to the bottom, may effectively influence fluid motion in the interior.
If we take AB = 10 cm2s1 and H = 100 m, we find a spin-down time T = 2H/(fB) 
6 days from (2.4.9) With A = 10 cm2s1 in the entire fluid, it will take about T ~ H2/A
~ 100 days before the friction in the interior effectively damps the motion.
The spin-down phenomenon also follows readily from Kelvin's circulation
theorem. In the interior of the fluid, where the effect of friction can be neglected, we
can apply (1.1.26) and (1.1.27), i.e.




G
v


r

k
abs

  (  vabs )   ( f   )  const.


(2.4.10)

where  is the area within the closed material curve  in the horizontal plane. The
vertical flux from the bottom Ekman layer must lead to horizontal motion in the
interior; see the sketch in Fig. 2.2.
65
Fig. 2.2. A material, closed curve at consecutive times t1 and t2.
In the present example, with a cyclone ( G  0) in the interior, leading to
convergence in the Ekman layer, a closed material curve in the interior horizontal
plane must be extended, as depicted in Fig. 2.2. Accordingly, the area  in (2.4.10)
becomes enlarged, and therefore the geostrophic vorticity must decrease, which
means spin-down of the cyclone.
We realize immediately that if we have an anti-cyclone ( G  0) in the interior,
this will result in a divergent mass transport in the bottom layer, and hence
contraction of material curves in the interior. According to (2.4.10), the absolute
vorticity then must increase. This increase is achieved by the (negative)  G
approaching zero, i.e. a spin-down of the anti-cyclone.
For a variable wind stress, a balance in the interior can be achieved if wB = wS.
From (2.4.2) and (2.4.3) we then obtain
G 

2 
k  (   S ) .
 f B
(2.4.11)
From our previous example with S /(L) ~ 106 cm s2, we find that this balance
results in  G ~ 4.5105s1.
2.5 Variable vertical eddy viscosity
Because of mixing due to the action of wind, breaking of waves etc., the level of
turbulence close to the surface is higher than in the interior of the ocean. This means
that the vertical eddy viscosity is not a constant, but depends on depth. Let us make a
simple study of how the effect of a vertical variation of A influences the Ekman flow
in the surface layer in a homogeneous ocean. For that purpose we consider a linear
depth dependence:
66
A  A0   z ,
(2.5.1)
where  can be positive or negative. In this way we model an A that is decreasing or
increasing with depth, respectively. We rescale the depth by
~
z  z / ,
0
(2.5.2)
where  0   (2 A0 / f )1 / 2 . Accordingly we can write
A  A0 (1   ~
z) .
(2.5.3)
Here  is a dimensionless parameter given by
 2
    
 A0 f
1/ 2



.
(2.5.4)
From (2.2.2), with p independent of x and y, we find that
(1   ~
z )W~z ~z  W~z  2 2iW  0 ,
(2.5.5)
where W = u + iv. We assume a constant wind stress in the y-direction. The boundary
conditions then become
W~z  i T , ~
z  0,
(2.5.6)
where T   S( y ) 0 /( A0 ) . In the deep ocean we assume that
W  0, ~
z   .
(2.5.7)
We will assume that  is a small parameter (|| << 1), allowing us to expand the
solution in a power series after . Accordingly
W  W ( 0)  W (1)   2W ( 2) 
(2.5.8)
We now insert this series into the governing equation (2.5.5) and the boundary
conditions (2.5.6) and (2.5.7). Then we collect terms with the same power of . Since
this is an arbitrary small parameter, our equations can only be fulfilled if the
coefficients multiplying each power of  vanish identically.
The terms multiplying 0 yield, when set to zero:
W~z(~z0 )  2 2iW ( 0 )  0,
~
W~( 0 )  iT ,
z  0,
z
(0)
W
 0,



~
z  . 
(2.5.9)
This is the ordinary Ekman problem for constant A, and the solution can be written
W ( 0)  C0 exp m~
z,
where m 2  2 2i and C0  iT / m .
(2.5.10)
67
Next, the terms multiplying 1 yield, when set to zero:
W~z(1~z)  2 2iW (1)   zW~z(~z0 )  W~z( 0 ) , 

~
W~z(1)  0,
z  0,

~

W (1)  0,
z  .

(2.5.11)
Here W(0) is given by (2.5.10). It is easy to show that a particular solution can be
written
Wp(1)  D~
z exp m~
z  E~
z 2 exp m~
z.
(2.5.12)
By insertion into (2.5.11) and comparing terms with equal powers of ~z , we find that
D  C0 / 4, E  mC0 / 4 .
(2.5.13)
A complete solution of (2.5.11) that satisfies W (1)  0, ~
z   , can then be written
W (1)  C1 exp m~
z  Wp(1) .
(2.5.14)
The application of the boundary condition W~z(1)  0, ~
z  0 , finally determines the
constant C1. The solution can be written
W (1) 
V0
(1  m~
z  m2 ~
z 2 ) exp m~
z,
4 2
(2.5.15)
where
V0 
 S( y )
 ( A0 f )1 / 2
.
(2.5.16)
In particular, let us study the surface current, i.e. we take ~
z  0 . The complex
surface flow W(0) can be written in the present approximation as
W (0)  W ( 0) (0)  W (1) (0) 
V0 
 
1  i 
.
4 
2
(2.5.17)
From the real and imaginary parts we obtain
V0 
 
1 
,
2  4 
V
v(0)  0 .
2
u (0) 
(2.5.18)
If we had continued our approximations, the next terms in (2.5.18) would have been
proportional to 2. If  is sufficiently small, these terms can be neglected. The surface
current is deflected an angle 0 to the right of the wind. This angle is given by
0  arctan
u (0)


 arctan 1 
v(0)
 4

.

(2.5.19)
68
We then observe that if  < 0 (A increases with depth), we have 0 < 45. For  > 0 (A
decreases with depth), we have 0 > 45. We must remember that this result has been
derived under the assumption that || << 1, or
1/ 2
A f 
   0 2 
 2 
.
(2.5.20)
Observations close to the surface often show a tendency for wind-driven ocean
currents (when geostrophic components are removed) to be more in the direction of
the wind than what follows from an Ekman analyses with constant eddy viscosity.
This can partly be due to the fact that some of the transport is caused by surface wind
waves in a direction close to the wind. However, it also suggests that the turbulent
eddy viscosity must increase with depth in the upper part of the surface layer, as
indicated by the present analysis. Since the eddy coefficients are small in the interior
of the ocean, the eddy viscosity must therefore increase to a maximum value at a
certain depth, before it decreases to attain its deep ocean value.
2.6 Equations for the volume transport
At the free surface z   ( x, y, t ) , we neglect any turbulent fluctuations. This
means that  can be associated with the mean value defined by (2.1.1). By integrating

the continuity equation   v  0 for the mean motion in the vertical, and applying the
mean versions of the boundary conditions (1.1.6) and (1.1.8), we find

t  



u dz 
v dz .

x  H
y H
(2.6.1)
This is identical to (1.1.10), but the dependent variables here are mean quantities.
When we take into account the effects of friction and a variable density, u and v are
no longer independent of z, as in section 1. We now introduce volume transports per
unit length. In the x- and y-direction they are, respectively




H


V   v dz. 

H
U
 u dz,
(2.6.2)
Accordingly, we can write (2.6.2) as
t  U x  V y .
(2.6.3)
69
In the momentum equation we apply the Boussinesq approximation, i.e. we
assume that the density changes are only important in connection with the action of
gravity. This means that in (2.1.11) we can use  = r, where r is a constant reference
density. Integrating the acceleration term in (2.1.11), we find






2
 H(ut  v  u)dz  U t  x Hu dz  y Huv dz .
(2.6.4)
From the assumption of hydrostatic balance in the vertical, we can write the
pressure as

p  PS ( x, y, t )  g   dz .
(2.6.5)
z
Let us now consider the x-component of (2.1.11). By integrating the pressure term
vertically, and applying the Boussinesq approximation, we find that


1
p
H
x
dz  
1
r

p
x
dz  
H
1
r
I x( p ) 
1
r
PS x 
1
r
PB H x .
(2.6.6)
Here PB and I(p) are the pressure at the bottom and the depth-integrated pressure,
respectively. They are defined by


PB  PS  g   dz, 

H


( p)

I   p dz.

H
(2.6.7)
The friction term containing the vertical variation becomes

 ( Au )
z z
dz 
H
1
r
( S( x )   B( x ) ) ,
(2.6.8)
where  S( x )   ( x ) ( ) is the wind stress and  B( x )   ( x ) (  H ) is the bottom stress in the
x-direction. The integrated friction terms containing the horizontal variation are more
difficult to write in a simple form. For example,

 2 

A
u
dz

A
u
dz


u
(

)

2

u
(

)

H
u
(

H
)

2
H
u
(

H
)

 , (2.6.9)
H
xx
H
xx
x
x
xx
x
x
2
 H

 x  H

where we have taken AH to be constant. We will follow Munk (1950), and assume that
the first term dominates on the right hand side, i.e.

A
H
H
u xx dz  AH U xx .
(2.6.10)
70
Quite similar results are obtained for the y-component of (2.1.11). Furthermore, we
assume that the integrated advection terms in (2.6.4) are small compared to the
integrated Coriolis term (i.e., we assume small Rossby numbers). The integrated
momentum equation can then be written
 r (U t  fV )   I x( p )  PS x  PB H x   S( x )   B( x )   r AH  2HU , 

 r (Vt  fU )   I y( p )  PS y  PB H y   S( y )   B( y )   r AH  2HV . 
(2.6.11)
These equations are very useful for describing oceanic transports. Even though we
have applied the Boussinesq approximation, they are valid for an ocean with variable
density, thus including baroclinic motion; see section 3.
In literature one often find these equations written in terms of mass transports
(M(x), M(y)) instead of volume transports, as here. The transition is very simple, since

M
( x)



   v dz   r  v dz   r V . 

H
H

M

  u dz  r  u dz   rU , 
H
( y)

H

(2.6.12)
An interesting challenge of great practical importance is to determine how the
surface elevation varies in time and space, when we know the atmospheric surface
pressure and the wind stresses from meteorological prognoses. For this particular
problem, it turns out that we to a good approximation can neglect the entire density
variation, i.e. treat the ocean as a constant density fluid. In that case we obtain from
(2.6.5):
p  PS ( x, y, t )  g r (  z ) .
(2.6.13)
Furthermore, we assume a simple, linear relation between the bottom stress and the
volume flux:
 B( x ) /  r  rU / H , 

 B( y ) /  r  rV / H , 
(2.6.14)
where r is a constant friction coefficient. This is analogous to (2.2.6) for the Ekman
current at the bottom, except that here we use the mean velocities U/H and V/H, and
not the bottom velocity. By assuming that   H , which always is a good
approximation, (2.6.11) reduces to
71
r

U, 
r
r
H


H
1 ( y) r
Vt  fU   gH y  PSy   S  V . 
r
r
H 
U t  fV   gH x 
H
PSx 
1
 S( x ) 
(2.6.15)
Here we have disregarded horizontal diffusion of momentum, i.e. taken AH = 0. These
equations, together with (2.6.3), govern the so-called storm surge problem for an
ocean of variable depth. The numerical solution of this system of equations, within
the appropriate geometrical domain, is performed routinely on a daily basis in many
countries bordering the sea to predict storm-induced water level changes along the
coasts.
We will not go into the general solution of this problem, but we mention that for
a storm area of limited extent, the solution will consist of a combination of free waves
“leaking” out of the area (propagating much faster than the storm centre) and a forced
solution like Ekman's elementary current system. This part of the solution has the
same lateral dimension as the storm area, and it follows approximately the motion of
the storm centre.
Equations (2.6.15) also describe how the motion becomes attenuated if the
atmospheric sea surface pressure becomes constant (PS = 0), and the wind ceases to

blow (  S  0 ). Let us assume that H is constant. By introducing the mean vorticity
 
1
(Vx  U y ) ,
H
(2.6.16)
we find on an f-plane that
t  
r
.
H
(2.6.17)
Accordingly, the integrated motion decreases exponentially with time. This is
analogous to what we found in (2.4.7) for the spin-down of a barotropic vortex due to
Ekman suction. We note that (2.6.17) is in fact correct even if PS  0, as long as the


wind stress has zero vorticity, i.e. if k  (  S )  0 .
As a second example of a transient solution of (2.6.15) for constant depth, we
consider planetary motion (  0). We now neglect the effect of bottom friction. For
an approximately stationary surface elevation (thereby excluding the presence of
Sverdrup waves), (2.6.3) reduces to
U x  Vy  0 .
(2.6.18)
72
We introduce a stream function  for the volume transport such that
U   y , 

V  x. 
(2.6.19)
By obtaining the vorticity from (2.6.15), and inserting from (2.6.19), we find


(2.6.20)
2Ht  x  k  (  S / r )
By comparing with (1.9.4), we realize that the homogenous solution of this equation
yields Rossby waves. We also note that the vorticity of the wind stress acts as a
source term for Rossby waves in the ocean.
2.7 The Sverdrup transport
For the rest of section 2 we are going to study ocean circulation. We simplify,
and assume that the surface pressure PS and the depth H are constants. Furthermore,
we consider stationary motion in the ocean. Then, in terms of the mass transports,
(2.6.3) becomes
M x( x)  M y( y )  0 ,
(2.7.1)
which naturally leads to the definition of a stream function  for the steady mass
transport:
M ( x )   y ,
M ( y)  x .
(2.7.2)
For this problem it proves convenient to derive an equation for the vertical vorticity
associated with the mass transports. With the present assumptions, and utilizing
(2.6.12), we find




 M ( y )  k  (  S )  k  (  B )  AH 2H ( M x( y )  M y( x ) ) .
(2.7.3)
By substituting the stream function from (2.7.2), we obtain a partial differential
equation for the single variable  :




 x  k  (  S )  k  (  B )  AH 4H .
(2.7.4)
Sverdrup (1947) assumed that the effects of friction could be neglected altogether. In
that case we need not utilize the stream function, since the mass transport in the northsouth direction M(y) is given directly by (2.7.3), i.e.


 M ( y )  k  (   S ) .
(2.7.5)
73
This relation is often referred to as the Sverdrup balance. From (2.7.1) we then obtain
that
M
( x)
x


  k  (   S ) 
 
.
y 


(2.7.6)
Sverdrup applied these equations to the equatorial region and assumed that the wind
(and the wind stress) was purely zonal, i.e.

 S( y )  0,

 S( x )   S( x ) ( y ). 
(2.7.7)
Then (2.7.6) for the east-west transport reduces to
M
( x)
x
1 d 2 S( x )
.

 dy 2
(2.7.8)
This equation is only 1. order in x, and accordingly we can only apply one boundary
condition. We take that M ( x )  0 at the eastern boundary. If this boundary is located
at x = L, the Sverdrup transports become, respectively:
M ( x) 
M
( y)
1 d 2 S( x )
( x  L) .
 dy 2
1 d S( x )
.

 dy
(2.7.9)
(2.7.10)
In Fig. 2.3 we have sketched the wind stress distribution in the equatorial region, and
the associated mass transport from (2.7.9).
Fig. 2.3. Sketch of the wind stress distribution and the mass transport in the
equatorial region.
74
Since M(x)  0 for x = 0, this solution is not valid close to the left (western) boundary.
We note from (2.7.9) that the equatorial transport is proportional to the 2. derivative
(the curvature) of the north-south distribution of the wind stress. By inserting an
equatorial easterly wind field (the trade winds) with a local minimum (the doldrums)
slightly north of the equator, Sverdrup demonstrated that we in principle obtain the
observed equatorial current system with westerly flowing currents north and south of
the equator and an easterly flowing counter current sandwiched between them; see the
sketch in Fig. 2.3. In reality, the equatorial flow pattern is more complicated, which is
probably due to the fact that the real wind system is not as simple as the one used
here.
An apparent paradox is revealed when we compare the Sverdrup east-west
transport, which is along the wind, to the Ekman transport, which is directed 90 to
the right of the wind. The solution to this problem lays in the fact that the Sverdrup
solution also contains the integrated pressure I(p), defined by (2.6.7). From the
stationary version of (2.6.11), with  S( y ) =  B( y ) =  = AH = 0 and H = const., we find
I y( p )   f M ( x ) .
(2.7.11)
f d S( x )
( x  L)  C ,
 dy
(2.7.12)
By inserting from (2.7.9):
I ( p)  
where C is a constant. In this expression, we note that x  L. Thus, we see that for a
given x, the integrated pressure I(p) has a maximum where d S( x ) / dy has its largest
value. This can only be a result of the transport to the right of the wind in the Ekman
layer close to the surface; see Fig. 2.4, where we have assumed a simple sinusoidal
wind profile.
75
Fig. 2.4. The connection between the wind direction, the Ekman transport and the
Sverdrup transport.
Accordingly, the Sverdrup transport in the east-west direction results from
geostrophic balance in the north-south direction.
Finally, we remark that the Sverdrup balance (2.7.5) expresses the fact that the
advection of planetary vorticity is equal to the vorticity of the surface wind stress. For


an anti-cyclonic wind field, as depicted in Fig. 2.4, we have that k  (  S ) =
 d S( x ) / dy  0 . Hence, a fluid column must move in such a way that it has a
southward-directed velocity component, as indicated in the figure.
2.8 Theories of Stommel and Munk (western intensification)
As emphasized in the previous section, Sverdrup’s solution for the east-west
transport does not satisfy the boundary condition at the western boundary. To remedy
this fault, the effect of friction must be taken into account. This was done by Stommel
(1948), and by Munk (1950).
We return to the more general equation (2.7.4). Stommel (1948) solved it with an
idealized wind stress distribution, while taking AH = 0. Munk (1950) neglected the

bottom friction, i.e. assumed  B  0 , and solved the equation using realistic wind
data. Munk's solution reproduces most of the features of the large-scale ocean
circulation. However, this solution is quite complicated. We will therefore be content
with Stommel’s simplified approach, which yields the main features of the circulation
76
pattern. However, we will return to Munk’s model when we discuss the boundary
layer at the western coast, using a simplified wind field.
Stommel modelled the bottom stress1 by assuming
 B( x ) K M ( x ) , 

 B( y ) K M ( y ) , 
(2.8.1)
where K is a constant friction coefficient. The ocean is taken to be a rectangular basin
of length L and width D, subject to a zonal wind field of the form:
 S( y )  0,
 S( x )


y 
  0 cos
.
D 
(2.8.2)
The ocean geometry and the wind stress are sketched in Fig. 2.5.
Fig. 2.5. Ocean basin and wind field for the Stommel solution.
With the parameterisations (2.8.1) and (2.8.2), and the assumption AH = 0, (2.7.4)
reduces to
K 2H   x  
 0
D
sin
y
D
.
(2.8.3)
The boundaries of the model are impermeable, which means that the flow must be
purely tangential at any part of the boundary. From the definition of the stream
function, this means that the boundary is a streamline given by  = 0.
We assume that the solution of (2.8.3) can be written
1
Stommel imagined that the lower limit of the vertical integration was taken to be the depth where the
current velocity was practically zero. This depth is usually much less than the total depth. In this case,
”the bottom stress” refers to the frictional stress between the dynamically active upper layer, and the
passive lower layer.
77
   ( x ) sin
y
D
.
(2.8.4)
This solution satisfies the boundary conditions for y =0, D, i.e. (0) = (D) = 0 for all
x. Inserting (2.8.4) into (2.8.3), we find
  

K
 
2
D
2
 
 0
KD
,
(2.8.5)
with boundary conditions   0, for x  0, L . This equation has a particular solution
p 
0D
,
K
(2.8.6)
while the homogeneous solution can be written
h  C1 exp( r1 x)  C2 exp( r2 x) .
(2.8.7)
Here
 2
2 
r1 , r2  



2K  4K 2 D 2 

1/ 2
.
(2.8.8)
The constants C1 and C2 are obtained from the boundary conditions. By insertion:
h (0)   p  0, 
h ( L)   p  0. 
(2.8.9)
 0D
{exp( r2 L)  1}

, 
 K {exp( r1L)  exp( r2 L)} 

 D
{exp( r1L)  1}
C2   0
.
 K {exp( r1L)  exp( r2 L)} 
(2.8.10)
This yields
C1 
The complete solution can then be written




exp{
(
L

x
)}
sinh(

x
)

exp{

x}sinh(  ( L  x)) 
 0D  y 
2K
2K

sin
1 
 , (2.8.11)
K
D 
sinh( L)



where
 2
2 
  2  2
D 
 4K
1/ 2
.
(2.8.12)
Let us first see what this solution yields on an f-plane, i.e. when  = 0. It is then
convenient to place the origin in the centre of the basin. This is done by defining new
coordinates:
x*  x  L / 2,
y*  y  D / 2 .
(2.8.13)
78
By insertion into (2.8.11), with  = 0, we find that

 L
 L

 
sinh{   x* }  sinh{   x* } 

 D
y
D 2
D 2

 .
 ( x* , y* )  0 cos * 1 

K
D 

sinh{ L}


D
(2.8.14)
We realize straightaway that the streamlines from (2.8.14) are symmetric about the
basin centre. This situation is depicted in Fig. 2.6.
Fig. 2.6. Stream lines when the beta-effect is zero.
This is not in accordance with our knowledge of the large-scale ocean circulation,
which tells us that the currents are much more intense on the western side (the Gulf
Stream, Kuroshio) than on the eastern side. We therefore conclude that the beta-effect
must be significant. By inserting relevant parameters ( ~ 21011 m1s1, K ~ 106s1,
D ~ 6103 km, L ~ 104 km) we find a flow pattern from (2.7.23) as sketched in Fig.
2.7.
Fig. 2.7. The flow pattern when the beta-effect is included.
79
In this case the solution clearly reproduces the intensified flow on the western side of
the ocean.
The northward transport on the western side of the ocean occurs mainly within a
thin boundary layer, where the thickness is determined by the effect of friction and the
beta-effect. For Stommel’s model, the typical boundary-layer scale S is
S ~ K /  .
(2.8.15)
The mass transport M b( y ) in the boundary layer is sketched in Fig. 2.8. We will return
to this problem at the end of this section.
Fig. 2.8. Sketch of the transport at the western boundary in Stommel’s model.
Finally, by a simple boundary-layer analysis we discuss the northward transport

at the western boundary in the Munk model (  B  0 , AH  0). By assuming that
 / x   / y in the boundary layer, the leading terms in (2.7.3) becomes
AH xxxx   x  0 .
(2.8.16)
Since M ( y )   x , we can insert M b( y ) directly into (2.8.16). Furthermore, the
boundary-layer scale M for this problem is defined by
 M  ( AH /  )1/ 3 .
(2.8.17)
By introducing a boundary-layer coordinate ~
x  x /  M , equation (2.8.16) can be
written
d 3 M b( y )
 M b( y )  0 .
3
~
dx
This equation has a solution of the form
M b( y )  C1 exp( r1~
x )  C2 exp( r2 ~
x )  C3 exp( r3 ~
x) ,
(2.8.18)
(2.8.19)
80
where
1 31 / 2
1 31 / 2
.
r1  1, r2    i
, r3    i
2
2
2
2
(2.8.20)
For the solution of (2.8.18) we must require that M b( y ) is finite when ~
x   , and that
M b( y )  0 when ~
x  0 (no-slip). Accordingly, C1 = 0 and C2 =  C3 in (2.8.19). The
solution can then be written as
 3 ~
M b( y )  C exp(  ~
x / 2) sin 
x  .
2


(2.8.21)
Here, the constant C must be determined by matching with the solution valid outside
the boundary layer, i.e. the Sverdrup transport. Utilizing the simplified zonal wind
stress (2.8.2), (2.7.10) yields
( y)
M Sv

1  0   y 
sin 
,  M  x  L.
 D
 D
(2.8.22)
In a closed basin, the total north-south transport must be zero. This constitutes the
required matching condition, i.e.

L
0
M
( y)
( y)
~
 M b  M dx   M Sv dx  0 .
(2.8.23)
Insertion from (2.8.21) and (2.8.22) allows us to determine C. When we utilize that M
<< L, the mass transport in the boundary layer becomes
M b( y ) 

2  0 L
x
exp  
3 M D
 2 M
  3x    y 
 sin 
 sin 
.

2

  M  D 
We have sketched this solution in Fig. 2.9.
Fig. 2.9. Sketch of the transport at the western boundary in the Munk model.
(2.8.24)
81
The dimensional thickness xb of the layer where the transport is northward, is given
by
xb 
2
3
M .
(2.8.25)
With AH ~ 103 m2s1 and  = 21011m1s1, we find that xb ~ 134 km. This fits well
with the width of the Gulf Stream.
A similar boundary-layer analysis for the Stommel model yields that
M b( y ) 
 0 L
 x
exp  
K D
 S
  y 
 sin 
,
  D 
(2.8.26)
where S = K/. If we choose a typical boundary-layer thickness xb    S , our
former values for the parameters of the problem gives that xb ~ 160 km.
It is worth noticing that in the strong shear in the western boundary current (see
 
 
Fig. 2.9), the Rossby number R ~ v  v /( f k  v ) is of the same order of magnitude
 

as the horizontal Ekman number E H ~ AH  2 v /( f k  v ) . This means that also the
nonlinear terms in the momentum equation may be important in this region.
3. BAROCLINIC MOTION
3.1 Two-layer model
We now proceed to study the effect of vertical density stratification in the ocean.
In many situations the density is approximately constant in a layer close to the
surface, while the density in the deeper water is also are constant (and larger). The
transition zone between the two layers is called the pycnocline. Thin pycoclines are
typically found in many Norwegian fjords. In extreme cases we can imagine that the
pycnocline thickness approaches zero, resulting in a two-layer model with a jump in
the density across the interface between the layers.
We start out by studying such a model. For simplicity we describe the motion in
reference system as shown in Fig. 3.1, where the x-axis is situated at the undisturbed
interface between the layers. The constant density in each layer is 1 and 2,
respectively, where 2 > 1.
82
Fig. 3.1. Model sketch of the two-layer system.
We assume hydrostatic pressure distribution in each layer. By applying that the
pressure is PS along the surface, and continuous at the interface, i.e. p1(z = ) = p2(z =
), we find that
p1   1 gz  1 g ( H 1 )  PS ,


p2   2 gz  g ( 2  1 )  1 g ( H 1 )  PS . 
(3.1.1)
We average the motion in the upper and lower layer:
(uˆ1 , vˆ1 ) 
1
h1
H 1 
 (u , v ) dz ,
1
1
(3.1.2)

1
(uˆ2 , vˆ2 ) 
(u2 , v2 ) dz ,
h2  H 2
(3.1.3)
Here, h1  H1     and h2    H2 are the total depths of the upper and lower
layers, respectively.
We assume that our equations can be linearized, i.e. we neglect the convective
accelerations. Furthermore, we will disregard the effect of the horizontal eddy
viscosity. This means that AH = 0 in each layer. By introducing volume transports
(U1 ,V1 )  h1 (uˆ1 , vˆ1 ), 

(U 2 ,V2 )  h2 (uˆ2 , vˆ2 ), 
the momentum equation for the upper layer becomes:
(3.1.4)
83

1
1
1

h1
1 ( y) 1 ( y) 
V1t  fU1   gh1 y  PS y   S   i . 

1
1
1
U1t  fV1   gh1 x 
h1
PS x 
1
 S( x ) 
1
 i( x ) , 
(3.1.5)
Here, (  S( x ) ,  S( y ) ) are the wind stresses along the surface, and (  i( x ) ,  i( y ) ) are the
internal frictional stresses between the layers.
Equivalently, for the lower layer we find
1
h
1
1

gh2 x  g*h2 x  2 PS x   i( x )   B( x ) , 
2
2
2
2

1
h2
1 ( y) 1 ( y) 
V2t  fU 2   gh2 y  g*h2 y 
P    B , 

2
2 S y 2 i
2
U 2t  fV2  
(3.1.6)
where (  B( x ) ,  B( y ) ) are the frictional stresses at the bottom. Furthermore, we have
defined
   1 
 g ,
g*   2
 2 
(3.1.7)
which is referred to as the reduced gravity, because the fraction ( 2  1 ) / 1 is small
for typical ocean conditions.
By integrating the continuity equation u x  v y  wz  0 in each layer, we find,
without any linearization of the boundary conditions, that
t  t  U1x  V1 y , 
t   U 2 x  V2 y . 
(3.1.8)
Barotropic response
Assume that the mean velocities in each layer are approximately equal, i.e.
uˆ1  uˆ2 , vˆ1  vˆ2 . This leads to
U1 U 2

,
H1 H 2
V1
V
 2 ,
H1 H 2
(3.1.9)
when we assume that  ,   H1, H 2 . For simplicity, we also take that the lower
layer has a constant depth. From (3.1.8) we then obtain
t   t  
or
H1
H
(U 2 x  V2 y )  1 t ,
H2
H2
(3.1.10)
84

H2
.
H1  H 2
(3.1.11)
Here the integration constant must be zero when we consider wave motion. We note
from (3.1.11) that  and  are in phase, and that | | < | |.
By neglecting the effect of the earth’s rotation, assuming constant surface
pressure and neglecting frictional effects in (3.1.5), equations (3.1.9) and (3.1.11)
yield
tt  g ( H1  H 2 ) xx  0 ,
(3.1.12)
when  / y  0 . The solution is
  F1 ( x  c0t )  F2 ( x  c0t ) ,
(3.1.13)
where c02  g ( H1  H 2 ) . The expression (3.1.13) describes long surface waves
propagating in a non-rotating canal with depth H1 + H2. This is the solution we would
have found if we, as a starting point, had neglected the density difference between the
layers; see the one-layer model in section 1. Such a solution (a free wave), which is
not influenced by the small density difference between the layers, is often referred to
as the barotropic response.
The original meaning of the word “barotropic” is related to the field of mass, and
expresses the fact that the pressure is constant along the density surfaces, i.e. the
isobaric and isopycnal surfaces coincide. Mathematically this can be expressed as
p    0 . This was the case for the free waves in section 1, where the density was
constant everywhere, and the pressure was constant along the sea surface. For the
two-layer model this would mean that the pressure should be constant along the
interface between the two layers. This is only approximately satisfied here, but
nonetheless it has become customary to denote the response in this case as the
barotropic response.
Baroclinic response
We now assume that t  t . Then, from (3.1.8):
(U 1  U 2 ) x  (V1  V2 ) y  0 .
(3.1.14)
For simplicity we take the bottom to be flat. A particular solution of (3.1.14) can be
written
85
U1  U 2 , 

V1  V2 , 
(3.1.15)
i.e. the volume fluxes are equal, but oppositely directed in each layer. By taking the
surface pressure to be constant, and neglecting the effect friction, summation of
(3.1.5) and (3.1.6) yields
 H   H 
    2 1 1 2  ,
 ( 2  1 ) H 2 
(3.1.16)
where, as in the barotropic case, the integration constant must be zero. Furthermore,
we have used that h1  H1, h2  H2. The difference between 1 and 2 is quite small,
which allows us to use the approximations 2 1 = , and 1 ~ 2  . Thus,
equation (3.1.16) can be rewritten as
 
  H1  H 2 
.
  H 2 
(3.1.17)
We note that  and  are oppositely directed, and that | | >> | |, as initially assumed.
Assuming that  / y  0 , we obtain from (3.1.6), (3.1.8), and (3.1.17) that
 tt  c12 xx  0 ,
(3.1.18)
where
c12  g *
H1 H 2
.
H1  H 2
(3.1.19)
The solution of (3.1.18) can be written
  F1 ( x  c1t )  F2 ( x  c1t ) .
(3.1.20)
This represents internal gravity waves propagating with phase speed c1 along the
interface between the layers; see the sketch in Fig. 3.2.
Fig. 3.2. Internal wave in a two-layer model.
86
The solution to (3.1.20) is often called the baroclinic response. As for barotropic, the
term “baroclinic” is linked to the mass field. In a baroclinic mass field the constant
pressure surfaces and the constant density surfaces intersect, i.e. p    0 .
Equation (3.1.17) shows that this is the case here, since when  > 0, then  < 0.
Accordingly, the pressure varies along the interface, which is a constant density
surface.
Let us assume that the lower layer is very deep, i.e. H2 >> H1. This is the most
common configuration in the ocean. From (3.1.6) we find for the x-component in the
lower layer

1
1
11
1
g x   g* x  PS x    i( x )   B( x )  U 2t  f V2  .
2
2
h2  2
2

(3.1.21)
For the baroclinic case, U2 and V2 are finite when h2  , and so are the frictional
stresses. Accordingly, for this limit, (3.1.21) reduces to
g x  
2
1
g* x  PS x .
1
1
(3.1.22)
In the same way we find for the y-component
g y  
2
1
g* y  PS y .
1
1
(3.1.23)
By inserting (3.1.22) and (3.1.23) into (3.1.5) for the upper layer, we find that
2
1
1

h1 g* x   S( x )   i( x ) , 
1
1
1



1
1
V1t  f U1  2 h1 g* y   S( y )   i( y ) . 
1
1
1

U1t  f V1 
(3.1.24)
For the baroclinic case, the depth of the upper layer can be written
h1  H1      H1   . Furthermore, we apply that 2  1   , and 1  2 = .
By linearizing the pressure term (the first term on the right-hand side), equations
(3.1.24) and (3.1.8) yield
U1t  f V1   g* H1h1x 
1
 S( x )   i( x ) , 


1

V1t  f U1   g* H1h1 y 
1
1


h1t  U1x  V1 y .
 ( y )   i( y ) , 
 S


(3.1.25)


These equations for the baroclinic response in the upper layer are formally identical to
the equations describing the storm surge problem for a homogeneous ocean; see
87
(2.6.15), when the upper layer thickness replaces the surface elevation, and the gravity
g is replaced by g * . The set of equations (3.1.25) describes what is often referred to as
a reduced gravity model for the volume transport in the upper layer. Even though the
numerical values for the volume fluxes in the lower layer are of the same order of
magnitude as in the upper layer, the mean velocity in the lower layer is negligible,
since H2  . Therefore, we usually say that the lower layer has no motion in this
approximation.
We immediately realize from (3.1.25) that transient phenomena such as
Sverdrup-, Kelvin- and planetary Rossby waves in a rotating ocean of constant
density have their internal (baroclinic) counter-parts in a two-layer model. The
analysis for the internal response is identical to the analysis is section 1. It often
suffices to replace g with g * and H with H1 in the solution for the barotropic
response.
Analogous to the barotropic case we can define a length scale a1 that
characterizes the significance of earth’s rotation. We write
a1  c1 / f ,
(3.1.26)
where c12  g * H 1 . The length scale a1 is called the internal, or baroclinic, Rossby
radius. Typical values for c1 in the ocean are 2-3 m s1. Hence, a1  20-30 km, which
is much less than the typical barotropic Rossby radius. Therefore, the effect of earth’s
rotation will be much more important for the baroclinic response than for the
barotropic one with the same horisontal scale, or wavelength.
3.2 Continuously stratified fluid
We now turn to the more general problem of a continuous density stratification,
and start by investigating the stability of a stratified incompressible fluid under the
influence of gravity. The equilibrium values are:

v  0,


  0 ( z ),

p  p0 ( z )   g  0 dz  const . 

(3.2.1)
We introduce small perturbations (denoted by primes) from the state of equilibrium,
writing the velocity, density, and pressure as
88
 
v  v ( x, y, z, t ),
  0 ( z )   ( x, y, z, t ),
p  p0 ( z )  p( x, y, z, t ).





(3.2.2)
We assume that the density is conserved for a fluid particle. Furthermore, we take that
the perturbations are so small that we can linearize our problem, i.e. neglect terms that
are products of perturbation quantities. The equations for the conservation of
momentum, density, and mass then reduce to
 0 ( z )(u t  fv)   p x   z( x ) ,
(3.2.3)
 0 ( z)(vt  fu)   p y   z( y ) ,
(3.2.4)
 0 ( z )wt   p z  g ,
t 
d0
w0,
dz
u x  v y  wz  0 .
(3.2.5)
(3.2.6)
(3.2.7)
Here we have for simplicity left out the primes that mark the perturbations. In (3.2.3)
and (3.2.4)  ( x ) and  ( y ) represent the tangential stresses in the fluid. We have also
assumed that the vertical variations of these stresses are much larger than the
corresponding horisontal variations. Furthermore, we have neglected the effect of
friction in vertical component of the momentum equation (3.2.5).
3.3 Free internal waves in a rotating ocean
We start by disregarding completely the effect of friction on the fluid motion, i.e.
we take  ( x )   ( y )  0 in (3.2.3) and (3.2.4). Furthermore, we introduce the BruntVäisälä frequency (or the buoyancy frequency) N, defined by
N 2 ( z)  
g d 0
.
 0 dz
(3.3.1)
We are here going to study motion in a stably stratified incompressible fluid. In this
case we must have that d0 / dz  0 , meaning that N is real and positive. Equation
(3.2.6) can then be written
g t  N 2  0 w  0 .
(3.3.2)
By differentiating (3.2.5) with respect to time, and utilizing (3.3.2), we find that
 0 (wt t  N 2 w)   p z t .
(3.3.3)
89
From this equation we note that the time scale for pure vertical motion ( pzt  0 ) is
N 1 . Elimination of the gradient from (3.2.3)-(3.2.4), yields the vorticity equation. On
an f-plane we obtain
(vx  u y )t  fwz ,
(3.3.4)
where we have applied (3.2.7). Forming the horizontal divergence from the same two
equations, and applying (3.2.7), we find
0 ( z){wzt  f (vx  u y )}  2H p .
(3.3.5)
Elimination of the vorticity from the equations above, yields
 0 wz t t  f 2 wz    2H pt .
(3.3.6)
Finally, by eliminating the pressure between (3.3.3) and (3.3.6), we obtain
 2

1
2 2
2 1
( w )  0 .
 H w  ( 0 wz ) z   N  H w  f
0
0 0 z z

t t
(3.3.7)
We simplify (3.3.7) by assuming that 0(z) varies slowly over the typical vertical scale
for w, i.e.
1
0
(  0 wz ) z  wz z .
(3.3.8)
This is Boussinesq approximation for internal waves. By introducing the BruntVäisälä frequency (3.3.1), we can write
N2
(  0 wz ) z  
wz  wz z .
0
g
1
(3.3.9)
We realize from (3.3.8) that the Boussinesq approximation implies that
N2
wz  wz z .
g
(3.3.10)
If d is a typical vertical scale for the motion, the above equation yields
N 2  g / d ,
(3.3.11)
where d max  H . For a shallow ocean we typically have that g/H ~ 101s2, while for
a deep ocean (H = 4000 m), the corresponding value becomes g / H ~ 14  102s2.
Measurements in the ocean show that N 2 ~ 104  106s2, so (3.3.11) is usually very
well satisfied. We will therefore utilize the Boussinesq approximation in the future
analysis of this problem. Equation (3.3.7) then reduces to
 2 wt t  N 2 ( z) 2H w  f 2 wz z  0 ,
(3.3.12)
90
where 2 is the three-dimensional Laplacian operator. We derive the same equation
by letting 0(z)  r on the left-hand sides of (3.2.3)-(3.2.5), where r is a constant
reference density. Then N 2  ( g /  r )d0 / dz . This latter approach is probably the
most common one when applying the Boussinesq approximation.
We assume that the ocean is unlimited in the horizontal direction, and consider a
wave solution of the form
w  W ( z ) exp( i (kx  t )) .
(3.3.13)
Here the x-axis is directed along the wave propagation direction. From (3.3.12) we
then obtain
 N 2  2 
W   k 2  2
W  0,
2 
  f 
(3.3.14)
where a prime denote differentiation with respect to z.
Constant Brunt-Väisälä frequency
Later on we shall allow N to vary with z. In this treatment, we simplify, and
assume that N is constant. Typical values for N and f in the ocean (and atmosphere)
are N ~ 102s1 and f ~ 104s1, i.e. N >> f. From equation (3.3.14) we then note that
we have wave solutions in the z-direction if f <  < N, while for  < f or  > N, the
solutions must be of exponential character in the z-direction.
Let us assume that f <  < N. We then take
W  exp( im z ) ,
(3.3.15)
where m is a wave number in the vertical direction. By insertion into (3.3.14), we
obtain the dispersion relation
2 
k 2 N 2  m2 f 2
.
k 2  m2
(3.3.16)
We realize that we have an-isotropic system, since  cannot be expressed solely as a
function of the magnitude of the wave number vector.
Analogous to the discussion of planetary Rossby waves in section 1.9, we can
define a wave number vector as

  ( k , m) .
Then the phase speed and group velocity, become, respectively


c   /  2 ,
(3.3.17)
(3.3.18)
91
and

c g    .
(3.3.19)
From the last relation we notice that the group velocity is always directed towards
increasing values of . When  is constant, we find from (3.3.16) that this gives isolines that are straight lines in the wave number space; see also the sketch in Fig. 3.3.
Fig. 3.3. Lines of constant frequency for internal waves with rotation in the twodimensional wave number space.
Along the m-axis (where k = 0), we have  = f (small). Along the k-axis (where m =
0), we have  = N (large). Hence, the direction of the group velocity becomes as
shown in the figure, while the phase speed from (3.3.18) is directed along the wave
number vector.
If we imagine that the wave number m is given, we can plot  as a function of k,
as depicted in Fig. 3.4.
Fig. 3.4. Dispersion diagram for internal waves with rotation.
92
We can define a Rossby radius of deformation for internal motion with vertical wave
number m by
ai 
N
.
mf
(3.3.20)
For k  ai1 , the effect of rotation dominates (compare with Fig. 1.6 for the
barotropic case).
In the ocean, the wave number m cannot be chosen arbitrarily, since the vertical
distance is limited by the depth. If we, for simplicity, disregard the surface elevation
and assume a constant depth, we must have that w = 0 for z = 0, H; see Fig. 3.5.
Fig. 3.5. Internal waves in an ocean with horizontal surface and horizontal bottom.
A solution of (3.3.14), which satisfies the upper boundary condition, is
W  C sin m z ,
(3.3.21)
where
 N 2  2 
.
m k  2
2 
  f 
2
2
(3.3.22)
For the solution to satisfy the boundary condition at z   H , we must require
m
n
,
H
n  1, 2, ...
(3.3.23)
This means that vertical wave number must form a discrete (but infinite) set. Equation
(3.3.22) then yields for the frequency
 k 2 N 2  f 2 n 2 2 / H 2 


2
2 2
2
 k n  /H

1/ 2
.
(3.3.24)
93
We see that, for a disturbance with a given wave number in the horizontal direction,
the system (ocean) responds with a discrete number of eigenfrequencies.
The solution for w in this case can be written
w  C sin( m z ) exp{i (kx  t )}
C
 exp{i (kx  mz  t )}  exp{i(kx  mz  t )}
2i
(3.3.25)
The latter expression can be interpreted as the superposition of waves in a horizontal
layer consisting of an incoming, obliquely upward propagating wave, and an
obliquely downward reflected wave, where m must attain the value (3.3.23) for the
wave system to satisfy the boundary condition at the bottom.
We now consider the case where the motion is mainly horizontal. This allows us
to disregard the vertical acceleration in the momentum equation, i.e. we apply the
hydrostatic approximation. Accordingly, in (3.3.3) we take that
wt t  N 2 w ,
(3.3.26)
which leads to
N 2w  
1
0
pz t .
(3.3.27)
From (3.3.26) we realize that the hydrostatic approximation leads to the condition
 2  N 2 .
(3.3.28)
From Fig. 3.4 we note that this implies that k << m, i.e. the horizontal scale of motion
is much larger than the vertical scale. Since the depth H yields the upper limit for the
vertical scale, disturbances with wavelength  >> H will satisfy the hydrostatic
condition. This requirement applies to barotropic surface waves as well as baroclinic
internal waves.
Applying the hydrostatic approximation, (3.3.16) reduces to

k2N 2 
  f 2 

m2 

1/ 2
.
(3.3.29)
This is the frequency for internal Sverdrup waves. For an ocean with depth H and a
horizontal surface, i.e. m = n/H as in equation (3.3.23), we can write
  ( f 2  c n2 k 2 )1 / 2 .
(3.3.30)
Here
cn  HN /( n ), n  1, 2, 3, ...
(3.3.31)
which is the phase speed for long internal waves in the non-rotating case. Since
94
m  n / H  N / cn ,
(3.3.32)
equation (3.3.20) yields the internal (baroclinic) Rossby radius
ai  an  cn / f , n  1, 2, 3, ...
(3.3.33)
We note that this is analogous to the definition of the barotropic Rossby radius
appearing in (1.3.14). For one single internal mode, i.e. a two-layer structure, this is
similar to (3.1.26).
3.4 Internal response to wind forcing; upwelling at a straight coast
We apply the set of equations (3.2.3)-(3.2.7), and utilize the Boussinesq
approximation and the hydrostatic approximation, i.e.
ut  fv  
vt  fu  
1
S
1
S

1
px 
 z( x ) , 
S

py 
1
S

( y)
z

,

p z   g .
(3.4.1)
(3.4.2)
Here the constant surface density S is used as a reference density. Furthermore, we
introduce the vertical displacement  ( x, y, z , t ) of a material surface, so that w = D/dt
in the fluid. Linearly, this becomes
w  t .
(3.4.3)
The conservation of density (3.2.6) then yields for the density perturbation

S
g
N 2 .
(3.4.4)
where we have assumed that  =  = 0 at t = 0. Inserting into (3.4.2), we obtain
p z    S N 2 ,
(3.4.5)
while the continuity equation can be written
 z t  u x  v y .
(3.4.6)
In general, we take that N = N(z), and we write the solutions to our problem as infinite
series. For simplicity, we assume that the depth is constant, and that the surface is
horisontal at all times. Accordingly:
  0, z  H , 0.
(3.4.7)
In principle, it is also possible to allow the surface to vary in time and space.
However, the solution shows that the internal response can be acchieved, to a good
95
approximation, by assuming a horisontal surface (the rigid lid approximation); see
Gill and Clark (1975). According to our adopted approach, we write the solutions as


n 1



v   vn ( x, y, t ) n ( z ),

n 1



pn   S  pn ( x, y, t ) n ( z ), 

n1


    n ( x, y, t ) n ( z ),

n 1


u   un ( x, y, t ) n ( z ),
(3.4.8)
where the primes denote differentiation with respect to z. By inserting the solutions
into (3.4.5), we find


n
 p ( x, y, t ) ( z)  p
n 1
n

n
n

N 2 ( z )  0 .

(3.4.9)
For the variables to separate, we must require
n
pn
 const. 
1
.
cn2
(3.4.10)
Furthermore, for (3.4.9) to be satisfied for all x, y and t, we must have that
n 
N2
n  0 .
cn2
(3.4.11)
The boundary conditions (3.4.7) yield
n  0, z   H , 0 .
(3.4.12)
Equation (3.4.11) and the boundary conditions (3.4.12) define an eigenvalue problem,
i.e. for given N = N(z) we can, in principle, determine the constant eigenvalues cn, and
the eigenfunctions n (z ) , which appear in the series (3.4.8).
It is easy to demonstrate that the differentiated eigenfunctions n constitute an
orthogonal set. Since (3.4.11) is valid for arbitrary numbers n and m, we can write
cn2n  N 2n  0, 

cm2 m  N 2m  0, 
(3.4.13)
where m  n. We multiply the upper and lower equations by m and n, respectively.
By subtracting and integrating from z = H to z = 0, utilizing (3.4.12), we find
0
(c  c )  nm dz  0 .
2
n
2
m
H
Accordingly, for n  m, i.e. cn  cm, we must have that
(3.4.14)
96
0
    dz  0,
n m,
n m
(3.4.15)
H
which proves the orthogonality. Since the eigenfunctions are known, apart from
multiplying constants (as for all homogeneous problems), we can normalize them by
assuming, for example, that
0

2
n
dz 
H
H
.
2
(3.4.16)
This procedure is generally valid for N = N(z). To exemplify, and discuss explicit
solutions in a simple way, we assume that N is constant. Then the eigenfunctions
become
N
 cn
n  An sin 

z  ,

(3.4.17)
which satisfies equation (3.4.11) and the upper boundary condition. The requirement
n(H) = 0 yields the eigenvalues:
HN
 n ,
cn
(3.4.18)
or cn  HN /( n ) , which is identical to (3.3.31). Finally, the normalization condition
(3.4.15) gives An = H/(n).
We now insert the series (3.4.8) into (3.4.1) and (3.4.6), and multiply each
equation with 1 , 2 , 3 , etc. By integrating from z = H to z = 0, and applying the
orthogonality condition (3.4.15), we finally obtain
un

 fvn  cn2 n   n( x ) ,
t
x
vn
2  n
 fun  cn
  n( y ) ,
t
y
 n
un vn


,
t
x
y








(3.4.19)
where

( x)
n

( y)
n
2

S H
2

S H

 ( x )


dz
,

n
 z

H

0
( y)

 H z n dz. 
0
(3.4.20)
We notice from (3.4.19) that this set of equations is formally identical to the equations
for the barotropic volume transports driven by surface winds, e.g. (2.6.15).
97
The problem in the baroclinic case is related to the form of the horizontal shear
stresses  ( x ) and  ( y ) in the fluid. In principle, these are unknown, and depending on
the fluid motion. However, we shall simplify the problem by assuming that we know
the vertical gradients of these stresses in the fluid.
Assume that a constant wind is blowing along a straight coast, so that the surface
wind stresses become  S( x )  0,  S( y ) = 0; see the sketch in Fig. 3.6. The model is
situated in the northern hemisphere, i.e. f > 0.
Fig. 3.6. Model sketch of upwelling/downwelling at a straight coast.
We assume that the shear stresses are only felt in a relatively thin layer close to the
surface, i.e. the mixed layer, with a thickness d << H. Here the stresses vary linearly
with depth:
 ( x)  z  d 

,
   S  d 

0,

( y)
  0,
( x)

 d  z  0, 

 H  z  d , 

 H  z  0. 

(3.4.21)
With this variation in z, (3.4.20) yields

2 S( x )
n (d ), 
 S Hd


 0.

 n( x )  

( y)
n
(3.4.22)
We assume that the solutions are independent of the along-shore coordinate x, i.e.,
from (3.4.19):
98
un
 fvn   n( x ) ,
t
vn

 fun  cn2 n ,
t
y
 n
vn

.
t
y








(3.4.23)
These equations have a particular solution where vn is independent of time. By
assuming that vn / t  0 , and eliminating un and n from the equations above, we
find
 2 vn
1
f
 2 v n  2  n( x ) ,
2
y
an
cn
(3.4.24)
where an  cn / f is the Rossby radius for internal waves. By requiring that
vn  0,
y  0, 

vn finite, y  ,
(3.4.25)
the solution of (3.4.25) becomes
vn  
 n( x )
f
1  exp(  y / an ) .
(3.4.26)
From (3.4.23) we then obtain
un   n( x )t exp(  y / an ),
 ( x )t
 n  n exp(  y / an ).
f an





(3.4.27)
Thus, un and n increase linearly in time during the action of the wind. From the
derived solution we see that a wind parallel to the coast results in upwelling or
downwelling within an area limited by the coast and the baroclinic Rossby radius.
Within this area we also notice the presence of a jet-like flow un parallel to the coast.
This flow is geostrophically balanced; see the second equation in (3.4.23) with
vn / t  0 .
We now discuss our solution in some more details. For this purpose the first term
in the series (3.4.8) for v and  suffice:
v  v11( z )  ... 

  11 ( z )  ... 
(3.4.28)
To simplify, we again take that N is constant. Then, from (3.4.17), (3.4.18) and
(3.4.22):
99

 z 
sin  ,


H 

c1  HN /  ,

( x)
2
 d  
 1( x )  S sin  . 
 S d  H  
1 
H
(3.4.29)
By inserting into (3.4.28), we find

2 S( x )
 d 
 z 
sin 
 1  exp(  y / a1 ) cos
  ... 
S f d  H 

H 

( x)
2 S
 d 
 z 
w  t 
sin 
 exp(  y / a1 ) sin 
  ..... 

S N d  H 
H


v
(3.4.30)
Here –H  z  0 and f > 0. For wind in the negative x-direction (  S( x ) < 0), we find that
w  0 in the region limited by baroclinic Rossby radius. Accordingly, the Ekman
surface layer transport away from the coast leads to a compensating flow from below
(upwelling). This is consistent with the sign of v in (3.4.30), since v is positive close
to the surface and negative near the bottom; see the sketch in Fig. 3.7.
Fig.3.7. Sketch of an upwelling situation.
We finally mention that since the x-component u and the vertical displacement 
increase linearly in time, the theory developed here is only valid as long as the
nonlinear terms in the equations remain small.
We return for a moment to the two-layer reduced gravity model to find out what
this would yield under similar conditions. By assuming  i( x ) =  i( y ) =  S( y ) = V1t = 0 in
(3.1.26), we find, analogous to (3.4.24):
V1 yy 
1
f  S( x )
V

,
1
a12
 c12
(3.4.31)
100
when we take that  / x  0 . The solution becomes
V1  

 S( x )
1  exp(  y / a1 ), 
f

 S( x )
t exp(  y / a1 ),

 ( x)
h1  S t exp(  y / a1 ),
 c1
U1 






(3.4.32)
where c1 = ( g * H1)1/2 and a1 = c1/f. We may define an upwelling velocity, when  S( x ) <
0, as
w1  h1t  
 S( x )
exp(  y / a1 ) .
c1
(3.4.33)
We can now compare with the case of continuous stratification. First, we assume that
the layer of frictional influence is thin, i.e. d << H. Furthermore, we insert for z =
H/2 to obtain the maximum vertical velocity. From (3.4.19) we then obtain
2 S( x )
w( z   H / 2)  
exp(  y / a1 ) .
  S c1
(3.4.34)
Here c1 = HN/ from (3.4.29). By comparing with (3.4.33), we see that the upwelling
velocities are remarkably similar, even though (3.4.34) is obtained from the first term
in a series expansion.
We will not go into further details of this problem. However, it is appropriate to
emphasize that this phenomenon is important for marine life. The water that upwells
is coming from depths below the mixed layer, and is rich in nutrients. Hence, the
upwelling process brings colder, nutrient-rich water to the euphotic zone, where there
is sufficient light to support growth and reproduction of plant algae (phytoplankton).
This means that upwelling areas are rich in biologic activity. Some of the world’s
largest catches of fish are made in such areas, e.g. off the coasts of Peru and Chile.
3.5 Baroclinic instability
When studying barotropic instability, we found that, under certain circumstances,
small perturbations in a homogeneous fluid could grow in time by extracting energy
from the kinetic energy of the basic flow. In a stratified, rotating fluid, where the
isobaric and the isopycnal surfaces are not parallel, there is another mechanism for
instability. In this case, small perturbations can grow by extracting energy from the
potential energy of the basic state. This instability is called baroclinic instability.
101
We consider the simple model described by Eady (1949). The geometry is
depicted in Fig. 3.8, and the isopycnals are assumed to be straight lines.
Fig. 3.8. Eady’s model.
The density 0 in the basic state can be written
 0   r (1   y   z ) ,
(3.5.1)
where ,  > 0 and r is a constant reference density. We assume that the basic state is
geostrophically balanced, and the vertical pressure distribution is hydrostatic.
Utilizing the Boussinesq approximation, the velocity- and pressure distributions can
be written
 g z 

v 0  Ui 
i,
f
p0  p r  g r ( z   yz  12  z 2 ) ,
(3.5.2)
(3.5.3)
respectively, where pr is a constant reference pressure. We introduce small
perturbations to this basic state, i.e. we write
 

v  v0 ( z )  v ( x, y, z, t ),


   0 ( y, z )   ( x, y, z, t ), 
p  p0 ( y, z )  p ( x, y, z, t ). 
(3.5.4)
Again, we assume that the (primed) perturbations are small enough for the equations
to be linearized. Furthermore, we disregard the effect of friction. The components of
the momentum equation can thus be written, respectively, as
ut  Uux  w
dU
1
 fv  
p ,
dz
r x
vt  Uv x  fu  
1
r
py ,
(3.5.5)
(3.5.6)
102
wt  Uwx  
1
r
pz 

g.
r
(3.5.7)
Here, for simplicity, we have skipped the primes associated with the perturbations.
Furthermore, we assume that the density is conserved for a particle, i.e. D / dt  0 .
Linearly, this means that
 t  U x  v
 0

 w 0  0.
y
z
(3.5.8)
The continuity equation then becomes
u x  v y  wz  0 .
(3.5.9)
We assume that f is constant. Equations (3.5.5) and (3.5.6) thus yield

 
g
  U
 
w y  f wz  0 ,
x 
f
 t
(3.5.10)
where   v x  u y . Here, we have also applied (3.5.2) and (3.5.9). In studying this
problem we assume that the development in time is slow, and that the velocities and
the velocity gradients are small. We may then assume that the horizontal velocity
perturbations are quasi-geostrophic, i.e.
1

py , 
f r


1
v
px . 
f r

u
(3.5.11)
Furthermore, we assume that the perturbation pressure is approximately hydrostatic in
the vertical:

1
pz .
g
(3.5.12)
When we apply the geostrophic assumption (3.5.11), the vorticity equation (3.5.10)
reduces to

 g

  2
  U
 H p  f r 
w y  fwz   0 .
x 
 f

 t
(3.5.13)
As far as the last two terms are concerned, we have that
g
wy
f
g H / f U max
~

.
fwz
fL
fL
(3.5.14)
103
In applying a quasi-geostrophic approximation, we have implicitly assume that the
Rossby number R0 = U/(fL) is much smaller than one. Thus, we can neglect the third
term in (3.5.13) compared to the fourth term. This equation then reduces to

  2
  U
 H p   r f 2 wz  0 .
x 
 t
(3.5.15)
Inserting for v and  from (3.5.11) and (3.5.12) into (3.5.8), yields

 
  U
 p z z  g r  wz  0 ,
 x 
 t
(3.5.16)
where we also have applied (3.5.1) and (3.5.2). Finally, by eliminating wz between
(3.5.15) and (3.5.16), we obtain a single equation for the perturbation pressure:


 
N2
  U
  p z z  2  2H p   0 .
 x 
f
 t

(3.5.17)
Here we have used that the Brunt-Väisälä frequency N can be written
g  0
 g .
 r z
N2  
(3.5.18)
We consider a wave solution in x and t. Then, in general, the differential operator
in (3.5.17) yields a non-zero factor. Accordingly, for (3.5.17) to be fulfilled for an
arbitrary wave component, we must have that
pz z 
N2 2
H p  0 .
f2
(3.5.19)
At the boundaries we assume that:
v  0,
w  0,
y  0, L, 

z  0, H . 
(3.5.20)
These are pure kinematic conditions. For this problem, they must be related to the
perturbation pressure at the boundaries, in order to solve (3.5.19). At the sidewalls, a
vanishing v in the geostrophic approximation (3.5.11) yields that
p  0,
y  0, L .
(3.5.21)
By inserting a vanishing w at the top and bottom boundaries into the density
conservation equation (3.5.8), we find:
 
g

p x  0,
  U  pz 
x 
f
 t
z  0, H .
Here we have also utilized (3.5.1), (3.5.11) and (3.5.12).
A wave solution that satisfies (3.5.21) can be written
(3.5.22)
104
p  P( z ) sin( ly ) exp{ik ( x  ct )} ,
(3.5.23)
where l  n / L for any integer n. We take that k is real, while c can be complex. By
inserting this solution into (3.5.19), we find
d2
P  4q 2 P  0 ,
dz 2
(3.5.24)
where
q
N 2 2 1/ 2
(k  l ) .
2f
(3.5.25)
The solution of (3.5.24) of can be written
P( z )  A sinh( 2qz )  B cosh( 2qz ) ,
(3.5.26)
where A and B are constants of integration. By applying (3.5.22) at z = 0, we obtain
that B  ( 2qcf / g ) A , i.e.


2qcf
P( z )  Asinh( 2qz ) 
cosh( 2qz )
g


(3.5.27)
By inserting (3.5.27) into (3.5.22) for z = H, we obtain an equation for the complex
phase speed c:
4q 2 f 2
g
2qH coth( 2qH )  1  0 .
c  4q 2 Hc 
g
f
(3.5.28)
After some algebra, we find that the solution to this equation can be written
c
U max
2

1
1/ 2 
1  q* coth q*  1q* tanh q*  1  ,
 q*

(3.5.29)
where q* = qH.
Let c  cr  ici in (3.5.23). Accordingly, if ci > 0, the perturbation solution will
grow exponentially in time, which means that the basic state is unstable. Oppositely,
the basic state is stable if ci < 0. Since always coth q*  1/ q* for q*  0 , it is obvious
from (3.5.29) that a root where ci > 0, requires that
q* tanh q *  1 i.e. q*  q*c  1.2 .
(3.5.30)
Inserting from (3.5.25), the condition for instability can be written
N 2H 2 2 2
(k  l )  4q*2c  5.76 .
f2
Unstable waves that satisfy (3.5.31), are usually referred to as Eady waves.
(3.5.31)
105
It is generally accepted that observed eddies in current systems like the Gulf
Stream and the Norwegian Coastal Current (NCC) may be caused by baroclinic
instability. Let us use NCC as an example. In a lecture note, Aas (1985) summarizes
some of the known facts about NCC. With reference to the parameters in the present
stability analysis, we can assume that H = 100 m and L = 40 km. Furthermore, we
take that the surface value of t is 23 close to the coast, and 27 where the coastal water
borders the Atlantic inflow. This yields  = 107m1 in equation (3.5.1). Insertion into
(3.5.2), with f = 1.3104s1, yields a surface current Umax= 0.76 m s1, i.e. a mean
northward flow of 0.38 m s1. Furthermore, if t varies from 23 to 27 over a vertical
distance of 100 m, then N = 2102s1 in (3.5.18). We consider the first mode (n = 1)
in the y-direction, which means that l = /L. We calculate the non-dimensional growth
rate kci / f from (3.5.29) as a function of the wavelength the  in the x-direction by
using the values assessed here. The result is depicted in Fig. 3.9.
Fig. 3.9 Non-dimensional growth rate as a function of the wavelength for Eady waves.
We observe from the figure that maximum growth occurs for waves with   80 km.
This scale fits well with observed length scales for meanders in NCC (Aas, 1985).
For the most unstable Eady wave with  = 80 km, (3.5.29) yields for the complex
phase speed in this example:
c  U max (0.5  0.17i ) .
(3.5.32)
106
The time Te it takes before the amplitude of the disturbance increases by a factor e
(the e-folding time) is given by
Te 
1
.
kci
(3.5.33)
The numerical value of the e-folding time for the most unstable wave in this example
is Te  27 hrs. The period for this wave, T  2 /( kcr ) , is here approximately 58 hrs.
We note that these time scales are sufficiently large to justify the application of the
quasi-geostrophic approximation, which requires that the time scale of the motion
should be considerably larger than the inertial period 2 /f (13 hrs in this example).
The growth rate of the disturbances calculated here is comparable with the observed
time scales for the development of meanders in NCC (Aas, 1985).
REFERENCES
Articles
Charney, J. G., and Eliassen, A.:
1949
Tellus, 1, 38.
Eady, E. A.:
1949
Tellus, 1, 33.
Eide, L. I.:
1979
Deep-Sea Res., 26, 601.
Gill, A. E., and Clarke, A. J.:
1974
Deep-Sea Res., 21, 325.
1979
J. Phys. Oceanogr., 9, 1126.
1981
Tellus, 33, 402.
Munk, W.:
1950
J. Met., 7, 79.
Pedlosky, J.:
1965
J. Mar. Res., 23, 207.
Stommel, H.:
1948
Trans. Am. Geophys. Un., 29, 202.
Sverdrup, H. U.:
1927
Geophys. Publ., 4, 75.
Sverdrup, H. U.:
1947
Proc. Nat. Acad. Sci., Wash., 33, 318.
Taylor, G. I.:
1916
Proc. Roy. Soc., Ser. A, XC 11, 196.
Taylor, G.I.:
1922
Proc. Roy. Soc., Ser. A, 102, 180.
Aas, E.:
1985
Lecture note, 27 pp.
Martinsen, E. A, Gjevik, B.,
and Røed, L. P.:
Martinsen, E. A.,
and Weber, J. E.:
107
Books
Defant, A.:
1961
Physical Oceanography, Vol. I & II. Pergamon Press 1961.
Gill, A. E.:
1982
Atmosphere-Ocean Dynamics. Academic Press 1982.
Krauss, W.:
1973
Methods and Results of Theoretical Oceanography. Vol. I
Gebrüder Borntraeger 1973.
LeBlond, P. H., and
Mysak, L. A.:
1978
Waves in the Ocean. Elsevier 1978.
Pedlosky, J.:
1987
Geophysical Fluid Dynamics, 2. ed. Springer 1987.
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