Casimir force at a knife's edge
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Graham, Noah et al. “Casimir force at a knife's edge.” Physical
Review D 81.6 (2010): 061701. © 2010 The American Physical
Society
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http://dx.doi.org/10.1103/PhysRevD.81.061701
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RAPID COMMUNICATIONS
PHYSICAL REVIEW D 81, 061701(R) (2010)
Casimir force at a knife’s edge
Noah Graham,1,* Alexander Shpunt,2 Thorsten Emig,2,3,4 Sahand Jamal Rahi,2 Robert L. Jaffe,2,5 and Mehran Kardar2
1
Department of Physics, Middlebury College, Middlebury, Vermont 05753, USA
Department of Physics, Massachusetts Institute of Technology, Cambridge, Massachusetts 02139, USA
3
Institut für Theoretische Physik, Universität zu Köln, Zülpicher Strasse 77, 50937 Köln, Germany
4
Laboratoire de Physique Théorique et Modèles Statistiques, CNRS UMR 8626, Bâtiment 100, Université Paris-Sud,
91405 Orsay cedex, France
5
Center for Theoretical Physics and Laboratory for Nuclear Science, Massachusetts Institute of Technology,
Cambridge, Massachusetts 02139, USA
(Received 24 October 2009; published 12 March 2010)
2
The Casimir force has been computed exactly for only a few simple geometries, such as infinite plates,
cylinders, and spheres. We show that a parabolic cylinder, for which analytic solutions to the Helmholtz
equation are available, is another case where such a calculation is possible. We compute the interaction
energy of a parabolic cylinder and an infinite plate (both perfect mirrors), as a function of their separation
and inclination, H and , and the cylinder’s parabolic radius R. As H=R ! 0, the proximity force
approximation becomes exact. The opposite limit of R=H ! 0 corresponds to a semi-infinite plate, where
the effects of edge and inclination can be probed.
DOI: 10.1103/PhysRevD.81.061701
PACS numbers: 03.70.+k, 12.20.m, 42.25.Fx
Casimir’s computation of the force between two parallel
metallic plates [1] originally inspired much theoretical
interest as a macroscopic manifestation of quantum fluctuations of the electromagnetic field in vacuum. Following
its experimental confirmation in the past decade [2], however, it is now an important force to reckon with in the
design of microelectromechanical systems [3]. Potential
practical applications have motivated the development of
numerical methods to compute Casimir forces for objects
of any shape [4]. The simplest and most commonly used
methods for dealing with complex shapes rely on pairwise
summations, as in the proximity force approximation
(PFA), which limits their applicability.
Recently we developed a formalism [5,6] that relates the
Casimir interaction among several objects to the scattering
of the electromagnetic field from the objects individually.
(For additional perspectives on the scattering formalism,
see the references in [6].) This approach simplifies the
problem, since scattering is a well-developed subject. In
particular, the availability of scattering formulas for simple
objects, such as spheres and cylinders, has enabled us to
compute the Casimir force between two spheres [5], a
sphere and a plate [7], multiple cylinders [8], etc. In this
work we show that parabolic cylinders provide another
example where the scattering amplitudes can be computed
exactly. We then use the exact results for scattering from
perfect mirrors to compute the Casimir force between a
parabolic cylinder and a plate. In the limiting case when the
radius of curvature at its tip vanishes, the parabolic cylinder becomes a semi-infinite plate (a knife’s edge), and we
can consider how the energy depends on the boundary
condition it imposes and the angle it makes to the plane.
The surface of a parabolic cylinder in Cartesian coordinates is described by y ¼ ðx2 R2 Þ=2R for all z, as
shown in Fig. 1, where R is the radius of curvature at the
tip. In parabolic cylinder coordinates [9], defined through
2
2
x ¼ , y p
¼ffiffiffið
ffi Þ=2, z ¼ z, the surface is simply
¼ 0 ¼ R for 1 < , z < 1. One advantage of
the latter coordinate system is that the Helmholtz equation
2
1
d d2 d2 r2 ¼ 2
þ
¼ 2 ; (1)
þ
þ 2 d2 d2
dz2
which we consider for imaginary wave number k ¼ i,
admits separable solutions. Since sending ! and
! returns us to the same point, we restrict our
attention to 0 while considering all values of .
*ngraham@middlebury.edu
1550-7998= 2010=81(6)=061701(5)
FIG. 1 (color online). Parabolic cylinder/plane geometry.
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PHYSICAL REVIEW D 81, 061701(R) (2010)
eikz z Z 1
i ðtan 2 Þ iky yþikx x
c out
dkx
e
;
ðrÞ ¼ pffiffiffiffiffiffiffi
ky cos2
8 1
Then plays the role of the ‘‘radial’’ coordinate in scattering theory and we have regular and outgoing wave
solutions
ik z
~
c reg
~
ðrÞ ¼ i e z D ðÞD ðiÞ;
out
ik
z
~ 1 ðÞ;
c ðrÞ ¼ e z D ðÞD
~
(2)
where D ðuÞ is the parabolic cylinder function, and ~ ¼
rffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffi
qffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffi
2 k2z þ 2 and similarly for . Enforcing the reflection
symmetry ! and ! for the regular solutions
restricts the separation constant to integer values. The
corresponding outgoing solutions do not obey this restriction and thus can only be used away from ¼ 0; as is
typical for outgoing solutions, they are irregular at ¼ 0.
For imaginary wave number, the regular (outgoing) solutions grow (decay) exponentially in and both i D ðiÞ
~
~ are real. We can then express the free scalar
and D ðÞ
Green’s function as [9]
1
1 Z 1 dkz X
1 reg
Gðr1 ; r2 ; Þ ¼ pffiffiffiffiffiffiffi
c ðr< Þ c out
ðr> Þ;
2
!
2 1
¼0
(3)
where r< (r> ) is the coordinate with the smaller (larger)
value of . We will also use the Green’s function in
coordinates appropriate to scattering from a plane perpendicular to the y axis,
Gðr1 ; r2 ; Þ ¼
Z 1 dkz
i
eikz ðz2 z1 Þ
4
2
1
Z 1 dkx
eikx ðx2 x1 Þþiky jy2 y1 j ;
1 ky
(4)
which is valid for y < 0.
The regular and outgoing waves provide two independent solutions to the second-order differential equation. We
take a linear combination of these solutions to obtain the
scattering solution ðrÞ outside the parabolic cylinder.
Fixing the coefficients by imposing Dirichlet boundary
conditions at ¼ 0 , we obtain
~ 0 Þ c reg
~ 0 Þ c out
ðrÞ ¼ D1 ð
ðrÞ i D ði
ðrÞ;
eikr
(7)
while for Neumann boundary conditions we have
þ1 0
~ 0 Þ c reg
D ði
~ 0 Þ c out
ðrÞ ¼ D01 ð
ðrÞ i
ðrÞ: (8)
These solutions to the Helmholtz equation can be used to
compute the Casimir forces between a parabolic cylinder
and other simple objects, for example, an infinite plate, as
depicted in Fig. 1. If both objects are perfect mirrors,
translational symmetry along the z axis enables us to
decompose the electromagnetic field into two scalar fields,
with Dirichlet and Neumann boundary conditions, respectively. Each scalar field can then be treated independently,
with the sum of their contributions giving the full electromagnetic result. The quantization of each scalar field is
achieved by integrating the exponentiated action over all
configurations of the field [10]. Constraining the fields to
obey the boundary conditions on each surface leads to an
alternative description involving fluctuating ‘‘charges’’
plane and cylinder on the surfaces [5,6]. Appropriate multipoles of these charges are
Z
dxdzdteikx xikz zþt plane ðx; z; tÞ;
QP ðkx ; kz ; Þ ¼
plane
Z
1
QC ðkz ; Þ ¼ pp
ddzdteikz zþt
ffiffiffiffiffiffiffiffiffiffiffiffiffiffi
ffiffiffiffiffiffiffi ffi
2! cylinder
2
2
c reg
ð; 0 Þcylinder ð; 0 ; z; tÞð þ 0 Þ:
qffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffi
where ky ¼ i 2 þ k2x þ k2z . We can connect the parabolic
and Cartesian Green’s functions using the expansion of a
plane wave in regular parabolic solutions [9]
1
X
1 ðtan2 Þ reg
¼
c ðrÞ;
¼0 ! cos 2
(6)
(9)
(5)
where tan ¼ kkxy . This expression converges in regions
where j tan2 j < 1. A plane wave with j tan2 j > 1 can
instead be expanded in terms of solutions with negative
integer values of [9], and the Green’s function can also be
expressed in terms of these functions analogously to
Eq. (3). Restricting to 0 is sufficient for our calculation, however, because we can already construct the
Green’s functions from these solutions alone; in the formalism of Refs. [5,6], all possible quantum fluctuations are
captured through the Green’s function. Equating Eqs. (3)
and (4) and then using (5), we obtain the expansion of the
outgoing parabolic solution in plane waves,
The action can be decomposed in terms of these multipoles as
Z1
Ldkz
S ¼
½S PP þ S CC þ S CP þ c:c:;
d
2
0
with
S PP ð; kz Þ ¼
i Z 1 dkx P
Q ðkx Þ ðF Pkx Þ1 QP ðkx Þ;
8 1 ky
1
1 X
QC ðF C Þ1 QC ;
(10)
2 ¼0 sffiffiffiffiffiffiffiffiffiffiffiffiffiffi
1 Z1
X
i
P
S CP ð; kz Þ ¼
dkx
Ukx ðd; ÞQC
Q ðkx Þ:
16k
y
¼0 1
S CC ð; kz Þ ¼ 061701-2
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Here S PP corresponds to the action for the charges on the
plane, with scattering amplitudes F Pkx ¼ 1 for Neumann
and Dirichlet modes, respectively. The corresponding action for charges on the parabolic cylinder S CC can be
related to its scattering amplitudes F C [6]; from Eqs. (7)
and (8) we obtain
D ði
~ 0Þ
ðDirichletÞ;
D1 ð
~ 0Þ
D0 ði
~ Þ
F C ¼ iþ1 0 0
ðNeumannÞ:
D1 ð
~ 0Þ
F C ¼ i
(11)
where is the angle of inclination of the parabolic cylinder
and d is the distance from the focus of the parabola to the
plane, as shown in Fig. 1.
Integrating over these charge fluctuations gives the
Casimir energy per unit length as
Z 1 d Z 1 dkz
E
¼
@cL
0 2 1 2
Z1
log det 10 F C
dkx Ukx ðd; ÞF Pkx
1
U0 kx ðd; Þ :
(13)
Numerical computations are performed by truncating the
determinant at index max . For the numbers quoted below,
we have computed for max up to 200 and then extrapolated
the result for max ! 1, and in the figures we have generally used max ¼ 100. We note that the integrals over and kz can be expressed as a single integral in polar
coordinates, and for ¼ 0 the kx integral is symmetric
and the translation matrix elements vanish for þ 0 odd.
Since the plane we are considering is a perfect mirror, F Pkx
is independent of kx and we can further simplify the
calculation for ¼ 0 using the integral
1
dkx
qffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffi
i ðtan2 Þ2n 2iky d
e
¼
2k
ð2d
2 þ k2z Þ;
2n1
ky cos2 2
(14)
where
1
10
100
1000
10 4
− 0.01
− 0.02
− 0.03
− 0.04
The position and orientation of the parabolic cylinder
relative to the plane enter only through the translation
matrix Ukx ðd; Þ, which appears in the interaction term
S CP . From Eq. (6), we obtain
sffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffi
ðtanþ
i
2 Þ
pffiffiffiffiffiffiffi
U kx ðd; Þ ¼
eiky d ;
(12)
2ky ! 2 cosþ
2
Z1
H/R
eu
‘
U ; 0; 2u
k ‘ ðuÞ ¼ ‘
2
ð2 þ 1Þ
is the Bateman k function [11], which is zero if ‘ is a
negative even integer. Here Uða; b; uÞ is the confluent
hypergeometric function of the second kind.
FIG. 2 (color online). The energy per unit length times H 2 ,
EH2 =ð@cLÞ, plotted versus H=R for ¼ 0 and R ¼ 1 on a loglinear scale. The dashed line gives the R ¼ 0 limit and the solid
curve gives the PFA result.
As a first demonstration, we report on the dependence of
the energy on the separation H ¼ d R=2 for ¼ 0. At
small separations (H=R 1) the PFA, given by
sffiffiffiffiffiffiffi
2 Z 1
dx
3
R
E PFA
pffiffiffi
;
¼
¼
2
3
720 1 ½H þ x =ð2RÞ
@cL
960 2 H5
(15)
should be valid. The numerical results in Fig. 2 confirm this
expectation with a ratio of actual to PFA energy of 0.9961
at H=R ¼ 0:25 (for R ¼ 1). We note that since the main
contribution to PFA is from the proximal parts of the two
surfaces, the PFA result in Eq. (15) also applies to a
circular cylinder with the same radius R.
A more interesting limit is obtained when R=H ! 0,
corresponding to a semi-infinite plate. Then the PFA result
is zero, as are results based on perturbative approximation
for the dilute limit [12]. The scattering amplitudes in
Eq. (11) simplify and can be combined together as F C ¼
pffiffiffiffiffiffiffiffiffi
! 2=, where even corresponds to Dirichlet and odd
corresponds to Neumann. Using this result, our expression for the energy for R ¼ 0 and ¼ 0 simplifies to
Z 1 qdq
E
C?
¼
log detð10 k0 1 ð2qHÞÞ ¼
;
@cL
H2
0 4
(16)
where C? ¼ 0:006 741 5 is obtained by numerical integration. This geometry was studied using the world-line
method for a scalar field with Dirichlet boundary conditions in Ref. [13]. The world-line approach requires a
large-scale numerical computation, and it is not known
how to extend this method to Neumann boundary conditions (or any case other than a scalar with Dirichlet boundary conditions). In our calculation, the Dirichlet
component of the electromagnetic field makes a contribu-
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PHYSICAL REVIEW D 81, 061701(R) (2010)
0.006
0.00684
0.00682
0.005
0.004
0.003
0.00680
0.00678
0.002
π/8
π/4
3π/8
π/2
0.00676
π/4
0.00674
π/8
3π/8
π/2
FIG. 3 (color online). The coefficient cðÞ as a function of
angle for R ¼ 0. The exact result at ¼ =2 is marked with a
cross. Inset: Dirichlet (circles) and Neumann (squares) contributions to the full electromagnetic result.
tion CD
? ¼ 0:006 048 5 to our result, in reasonable agreement with the value of CD
? ¼ 0:006 00ð2Þ in Ref. [13].
Reference [13] also considers a tilted semi-infinite plate,
which corresponds to the R ! 0 limit of our formula for
general . From dimensional analysis, the Casimir energy
at R ¼ 0 again takes the now -dependent form
CðÞ
E
¼ 2 :
H
@cL
(17)
Following Ref. [13], we plot cðÞ ¼ cosðÞCðÞ in Fig. 3.
A particularly interesting limit is ! =2, when the two
plates are parallel. In this case, the leading contribution to
the Casimir energy should be proportional to the area of the
half-plane according to the parallel plate formula,
Ek =ð@cAÞ ¼ ck =H 3 with ck ¼ 2 =720, plus a subleading correction due to the edge. Multiplying by cos removes the divergence in CðÞ as ! =2. As in Ref. [13],
we assume cð ! =2Þ ¼ ck =2 þ ð =2Þcedge , although we cannot rule out the possibility of additional
nonanalytic forms, such as logarithmic or other singularities. With this assumption, we can estimate the edge
correction cedge ¼ 0:0009 from the data in Fig. 3. From
the inset in Fig. 3, we estimate the Dirichlet and Neumann
contributions to this result to be cD
edge ¼ 0:0025 (in
agreement with [13] within our error estimates) and
cN
edge ¼ 0:0034, respectively. Because higher partial waves
become more important as ! =2, reflecting the divergence in CðÞ in this limit, we have used larger values of
max for near =2.
It is straightforward to extend these resultsRto nonzero
d
temperature T. We simply replace the integral 1
0 2 by the
T P1
0
over Matsubara frequencies n ¼
sum @c n¼0
2nT=ð@cÞ, where the prime indicates that the n ¼ 0
mode is counted with a weight of 1=2 [6]. In the limit of
infinite temperature, only the n ¼ 0 mode contributes and
we obtain for R ¼ 0 the energy E=L ¼ TCT¼1 =H, with
CT¼1 ¼ 0:0472. The Dirichlet contribution to our result is
CD
T¼1 ¼ 0:0394, again in agreement with [13].
Employing the scattering formalism, we can also calculate the Casimir energy for the case where another object
whose scattering amplitudes are available, such as an
ordinary cylinder or a second parabolic cylinder, is positioned outside the parabolic cylinder. Centering the other
object at the origin and letting the parabolic cylinder open
downward, with its focus displaced to y ¼ d, we obtain
the necessary translation matrix elements by writing
where x ¼ x, y ¼ y d, z ¼ z, and then
Eq. (6) for r,
expanding the plane wave on the right-hand side in the
basis appropriate to the other object. Again we can allow
the parabolic cylinder to tilt by replacing by þ in
this expression. These results can be extended to multiple
objects, as in Ref. [14]. Another interesting possibility
would be to apply the interior Casimir formalism of
Ref. [15] an object inside a parabolic cylinder, potentially
extending the results of Refs. [16,17].
The reduction of the parabolic cylinder to a semi-infinite
plate enables us to consider a variety of edge geometries. A
thin metal disk perpendicular to a nearby metal surface
would experience a Casimir force described by an extension of Eq. (16). Figure 2 shows that the PFA breaks down
for a thin plate perpendicular to a plane; the PFA approximation to the energy vanishes as the thickness goes to zero,
while the correct result instead has a different power law
dependence on the separation. Based on the full result for
perpendicular planes, however, we can formulate an ‘‘edge
PFA’’ that yields the energy by integrating dE=dL from
Eq. (16) along the edge of the disk. Letting r be the disk
radius, in this approximation we have
pffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffi
Zr
E EPFA ¼ @cC?
ðH þ r r2 x2 Þ2 dx
r
qffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffiffi
H=r!0
! @cC? r=ð2H 3 Þ;
which is valid if the thickness of the disk is small compared
to its separation from the plane. (For comparison, note that
the ordinary PFA for a metal sphere of radius r and a plate
is proportional to r=H 2 .)
A disk may be more experimentally tractable than a
plane, since its edge does not need to be maintained
parallel to the plate. One possibility would be a metal
film, evaporated onto a substrate that either has low permittivity or can be etched away beneath the edge of the
deposited film. Micromechanical torsion oscillators, which
have already been used for Casimir experiments [18], seem
readily adaptable for testing Eq. (17). Because the overall
strength of the Casimir effect is weaker for a disk than for a
sphere, observing Casimir forces in this geometry will
require greater sensitivities or shorter separation distances
than the sphere-plane case. As the separation gets smaller,
however, the dominant contributions arise from higherfrequency fluctuations, and deviations from the perfect
conductor limit can become important. While the effects
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of finite conductivity could be captured by an extension of
our method, the calculation becomes significantly more
difficult in this case because the matrix of scattering amplitudes is no longer diagonal.
To estimate the range of important fluctuation frequencies, we consider R H and ¼ 0. In this case, the
integrand in Eq. (16) is strongly peaked around q
0:3=H. As a result, by including only values of q up to
2=H, we still capture 95% of the full result (and by going
up to 3=H we include 99%). This truncation corresponds to
a minimum fluctuation wavelength min ¼ H. For the
perfect conductor approximation to hold, min must be
large compared to the metal’s plasma wavelength p , so
that these fluctuations are well described by assuming
perfect reflectivity. We also need the thickness of the disk
to be small enough compared to H that the deviation from
the proximity force calculation is evident (see Fig. 2), but
large enough compared to the metal’s skin depth that the
perfect conductor approximation is valid. For a typical
metal film, p 130 nm and
25 nm at the relevant
wavelengths. For a disk of radius r ¼ 100 m, the present
experimental frontier of 0.1 pN sensitivity corresponds to a
separation distance H 350 nm, which then falls within
the expected range of validity of our calculation according
to these criteria. The force could also be enhanced by
connecting several identical but well-separated disks. In
that case, the same force could be measured at a larger
separation distance, where our calculation is more
accurate.
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We thank U. Mohideen for helpful discussions. This
work was supported by the National Science Foundation
(NSF) through Grants No. PHY05-55338 and No. PHY0855426 (N. G.), No. DMR-08-03315 (S. J. R. and M. K.),
Defense Advanced Research Projects Agency (DARPA)
Contract No. S-000354 (S. J. R., M. K., and T. E.), by the
Deutsche Forschungsgemeinschaft (DFG) through Grant
No. EM70/3 (T. E.), and by the U. S. Department of Energy
(DOE) under cooperative research agreement No. DFFC02-94ER40818 (R. L. J.).
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