Universality for Shape Dependence of Casimir Effects from Weyl Anomaly

# Universality for Shape Dependence of Casimir Effects from Weyl Anomaly

Rong-Xin Miao,111Corresponding author. Physics Division, National Center for Theoretical Sciences,
National Tsing-Hua University, Hsinchu 30013, Taiwan Department of Physics, National Tsing-Hua University, Hsinchu 30013, Taiwan
Chong-Sun Chu, Physics Division, National Center for Theoretical Sciences,
National Tsing-Hua University, Hsinchu 30013, Taiwan Department of Physics, National Tsing-Hua University, Hsinchu 30013, Taiwan
###### Abstract

We reveal elegant relations between the shape dependence of the Casimir effects and Weyl anomaly in boundary conformal field theories (BCFT). We show that for any BCFT which has a description in terms of an effective action, the near boundary divergent behavior of the renormalized stress tensor is completely determined by the central charges of the theory. These relations are verified by free BCFTs. We also test them with holographic models of BCFT and find exact agreement. We propose that these relations between Casimir coefficients and central charges hold for any BCFT. With the holographic models, we reproduce not only the precise form of the near boundary divergent behavior of the stress tensor, but also the surface counter term that is needed to make the total energy finite. As they are proportional to the central charges, the near boundary divergence of the stress tensor must be physical and cannot be dropped by further artificial renormalization. Our results thus provide affirmative support on the physical nature of the divergent energy density near the boundary, whose reality has been a long-standing controversy in the literature.

## 1 Introduction

The Casimir effect Casimir:1948dh () originates from the effect of boundary on the zero point energy-momentum of quantized fields in a system. As a fundamental property of the quantum vacuum, it has important consequences on the system of concern and has been applied to a wide range of physical problems, such as classic applications in the study of the Casimir force between conducting plates (and nano devices) Plunien:1986ca (); Bordag:2001qi (), dynamical compactification of extra dimensions in string theory App1 (); App2 (), candidate of cosmological constant and dark energy Milton:2004ya (), as well as dynamical Casimir effect and its applications dyn ().

The near boundary behavior of the stress tensor of a system is crucial to the understanding of the Casimir effect. For a Quantum Field Theory (QFT) on a manifold of integer dimension and boundary , the renormalized stress tensor is divergent near the boundary Deutsch:1978sc ():

 ⟨Tij⟩=x−dT(d)ij...+x−1T(1)ij,x∼0, (1)

where is the proper distance from the boundary and with depend only on the shape of the boundary and the kind of QFT under consideration. For CFT with conformal invariant boundary condition (BCFT), one further require that divergent parts of renormalized stress tensor are traceless in order to get a well-defined finite Weyl anomaly without divergence. It is also natural to impose the conservation condition of energy:

 limx→0⟨Ti i⟩=O(1),∇i⟨Ti j⟩=0. (2)

Substituting (1) into the above equations, Deutsch:1978sc () obtains

 T(d)ij=0, T(d−1)ij=2α1¯kij, (3a) T(d−2)ij=−4α1d−1n(ihlj)∇lk−4α1d−2n(ihlj)npRlp +2α1d−2(ninj−hijd−1)Tr¯k2+tij, (3b)
 tij:=⌈β1Cikjlnknl+β2Rij+β3kkij+β4kliklj⌉, (4)

where , and are respectively the normal vector, induced metric and the traceless part of extrinsic curvature of the boundary . The tensor is tangential: , denotes the traceless part, is Weyl tensor of and is the intrinsic Ricci tensor of . The coefficients fixes the shape dependence of the leading and subleading Casimir effects of BCFT. The main goal of this letter is to show that one can fix completely these Casimir coefficients in terms of the bulk and boundary central charges.

## 2 Shape Dependence of Casimir effects from Weyl Anomaly

Consider a BCFT with a well defined effective action. The Weyl anomaly , defined as the trace of renormalized stress tensor, can be obtained as the logarithmic UV divergent term of the effective action,

 I=⋯+Alog(1ϵ)+Ifinite, (5)

where denotes terms which are UV divergent in powers of the UV cutoff , and is the renormalized, UV finite part of the effective action. This part is dependent on the subtraction scheme. But the dependence is irrelevant for the discussion below and our results hold for any renormalization scheme.

Inspired by Lewkowycz:2014jia (); Dong:2016wcf (), let us regulate the effective action by excluding from its volume integration a small strip of geodesic distance from the boundary. Then there is no explicit boundary divergences in this form of the effective action, however there are boundary divergences implicit in the bulk effective action which is integrated up to distance . The variation of effective action is given by

 δI=12∫x≥ϵ√g^Tijδgij (6)

where is the non-renormalized bulk stress tensor. The renormalized bulk stress tensor is defined by the difference of the non-renormalized bulk stress tensor against a reference one Deutsch:1978sc ():

 Tij=^Tij−^Tij0, (7)

where is the non-renormalized stress tensor defined for the same CFT without boundary. It is

 δI0=12∫x≥ϵ√g^Tij0δgij, (8)

where is the effective action of the CFT with the boundary removed, hence the integration over the region . Subtract (8) from (6) and focus on only the logarithmically divergent terms, we obtain our key formula

 (δA)∂M=(12∫x≥ϵ√gTijδgij)log(1/ϵ), (9)

where is the boundary terms in the variations of Weyl anomaly and is the renormalized bulk stress tensor. In the above derivations, we have used the fact that and have the same bulk Weyl anomaly so that

 (δA)∂M=(δI−δI0)log(1/ϵ). (10)

We observe that as the right hand side of (9) must give an exact variation, this imposes strong constraints on the possible form of the stress tensor near the boundary since this is where one would pick up logarithmic divergent contribution on integration near the boundary. It is this integrability of the variations which helps us to fix the Casimir effects in terms of the Weyl anomaly. To proceed, let us start with the metric written in the Gauss normal coordinates

where . The coefficients , , parametrize the derivative expansion (with respect to both and ) of the metric. Consider variation of the metric with and . Take first the 3d BCFT as an example. The Weyl anomaly of 3d BCFT is given by Jensen:2015swa ()

 A=∫P√h(b1R+b2Tr¯k2), (12)

where are boundary central charges which depends on the boundary conditions. Taking the variation of (12), we have

 b2∫P√h[(Tr¯k22hab−2¯kackcb)δhab+2¯kabδkab]. (13)

Now we turn to calculate the variation of Weyl anomaly from the last term of (9). Note that for . Note also that , where is the bivector of parallel transport between and Deutsch:1978sc (). Taking these facts into account and substitute (1) and (3) into the last term of (9), integrate over and select the logarithmic divergent term, we obtain

 − α1∫P√h[(Tr¯k22hab−2¯kackcb)δhab+2¯kabδkab] (14) + ∫P√h[(β32−α1)k¯kabδhab+β42⌈kackcb⌉δhab].

Note that (14) is made up of a structure of curvature components different from those appearing in (13). Integrability of (14) gives and . Comparing (13) with (14) gives . All together, we obtain the relations between the Casimir coefficients of the stress tensor and the boundary central charges:

 α1=−b2,β3=−2b2,β4=0. (15)

Similarly for 4d BCFT, we can obtain the shape dependence of Casimir effects from the Weyl anomaly Fursaev:2015wpa (); Herzog:2015ioa ()

 A = ∫M√g(c16π2CijklCijkl−a16π2E4) (16) +∫P√h(b3Tr¯k3+b4Cac   bc¯kb a),

where are bulk central charges and are boundary central charges. is the Euler density including the boundary term. To derive , we set for simplicity, since it only affects the third order derivative terms in the stress tensor. Taking variation of (16) and comparing the boundary term with the last term of (9), we obtain

 α1=b42,β1=c2π2+b4,β2=0,β3=2b3+136b4,β4=−3b3−2b4. (17)

It is remarkable that the boundary behavior of the stress tensor is completely determined by the boundary and bulk central charges However, it is independent of the central charge related to Euler density due to the fact that topological invariants do not change under local variations. We propose that the relations (15) and (17) between Casimir coefficients and central charges hold for general BCFT.

## 3 Free and Holographic BCFT

Let us verify our general statements with free BCFT. The renormalized stress tensor of 4d free BCFT has been calculated in Deutsch:1978sc (); Kennedy:1979ar (); Kennedy:1981yi (). The bulk and boundary central charges for 4d free BCFTs were obtained in Fursaev:2015wpa (). We summary these results in Table 1 and Table 2. Note that the results for Maxwell field apply to both absolute and relative B.C. We find these data obey exactly the relations (17). for Maxwell field is absence in the literature. Here from (17), we predict that for all 4d free BCFT due to the fact that for 4d free BCFT. As we will show below, this relation is violated by strongly-coupled CFT dual to gravity. As a result, is non-zero in general. Comparing with Kennedy:1981yi (), we note that there is a minus sign typo of for Maxwell field in Deutsch:1978sc ().

Now let us investigate the shape dependence of Casimir effects in holographic models of BCFT. Consider a BCFT defined on a manifold with a boundary . Takayanagi Takayanagi:2011zk () proposed to extend the dimensional manifold to a dimensional asymptotically AdS space so that , where is a dimensional manifold which satisfies . The gravitational action for holographic BCFT is Takayanagi:2011zk () ()

 I = ∫N√G(R−2Λ)+2∫Q√γ(K−T) (18)

plus terms on and . Here is a constant which can be regarded as the holographic dual of boundary conditions of BCFT Miao:2017gyt (); Chu:2017aab (). A central issue in the construction of the AdS/BCFT is the determination of the location of in the bulk. Takayanagi:2011zk () propose to use the Neumann boundary condition

 Kαβ−(K−T)γαβ=0 (19)

to fix the position of . In Miao:2017gyt (); Chu:2017aab () we found there is generally no solution to (19) for bulk metric that arose from the FG expansion of a general non-symmetric boundary. The reason is because is of co-dimension one and we only need one condition to determine it’s position, while there are too many extra conditions in (19). To resolve this, we suggested in Miao:2017gyt (); Chu:2017aab () to use the trace of (19), , to determine the position of . Nonetheless, it is also possible that one may need to relax the assumption that the bulk metric admits a valid FG expansion, as has been attempted in Nozaki:2012qd () for some non-symmetric boundary in BCFT. In contrast to a FG-expanded metric whose form near the boundary is completely fixed, a non-FG expanded metric has more degree of freedom. It was suggested in Nozaki:2012qd () that the embedding equation (19) may admit a solution if the bulk metric is also allowed to adjust itself. However in general this is a highly non-trivial problem and there is no systematic method available to construct gravity solutions for BCFT in general dimensions and with an arbitrary non-symmetric boundary () that is not FG expanded. Remarkably this problem can solved and we will now present the solution.

To make progress in this front, we find that one can instead consider an expansion in powers of small derivatives of the metric and keep both the and dependence as exact to construct a perturbative solution to the Einstein equation. For simplicity, we consider the case of here. The more general case of a nontrivial boundary metric can be analysed. We comment on this in the supplementary information. We find useful to consider the following metric ansatz

with a function such that . To find solution, let us first consider the region and consider the ansatz . This ansatz plays an important role to solve (19) for non-symmetric boundary with . For simplicity we consider a traceless extrinsic curvature here. The solution for the general case is given in the supplementary information. Substituting (20) into Einstein equation and writing , we obtain at the order a single equation

 s(s2+1)f′′(s)−(d−1)f′(s)=0. (21)

It has the solution

 f(s)=1−α1sd2F1(d−12,d2;d+22;−s2)d. (22)

To obtain a solution of the Einstein equation for , one may analytic continuate (22) to the region . However this solution while continuous at , is discontinuous at as the region near is mapped to widely separated regions . Another possibility is to first rewrite the expression (22) in terms of and , and then analytic continuate the resulting function to the region . In this way, we obtain a solution of the Einstein equation that is continuous at . For example, for , we have

 f(x,z)=1−α1(zx−g(x,z)), (23a) g(x,z)=π2−2tan−1(x/(z+√z2+x2)). (23b)

Let us make some comments. 1. For general , the perturbation is finite which shows that (21) is a well-defined metric. 2. Note that formally one can expand as a power series of and interpret that as a FG expansion of the metric (20). However the series does not converge whenever . Therefore for the boundary () physics we are interested in, it is necessary to use the exact solution without performing the FG expansion. 3. The perturbative background (20), (22) to the Einstein equation is an interesting result which may be useful for other studies as well.

So far the coefficient is arbitrary. If we now consider (19) in this background, we find that one can solve the embedding function of as provided that is fixed at the same time. Please see the supplementary information for more details. See Table 3 for values of obtained from holography, where we have re-parametrized and is the angle between and the bulk boundary . Using (20), (22), we can derive the holographic stress tensor deHaro:2000vlm ()

 Tij=limz→0dδgijzd=2α1¯kijxd−1+O(k2), (24)

which takes the expected form (3a). According to deHaro:2000vlm (), (24) automatically satisfy the traceless and divergenceless conditions (2). Note that in general the stress tensor (24) also contains contributions from in even dimensions deHaro:2000vlm (). However, these contributions are finite, so we can ignore them without loss of generality since we focus on only the divergent parts in this letter.

Similarly, we can work out the next order solutions to both the Einstein equation and (19), and then derive the stress tensor up to the order by applying the formula (24). See the appendix for details. It turns out that the holographic stress tensor takes exactly the expected expression (3) with the coefficients listed in Table 3.

These coefficients indeed satisfy the relations (15), (17) provided the boundary central charges are given by note ()

 b2=1θ, (25a) b3=11+tanhρ−13,b4=−11+tanhρ, (25b)

for 3d and 4d respectively. Since we have many more relations (8) than unknown variables (3), this is a non-trivial check of the universal relations (15), (17) as well as for the holographic proposal (19). In fact, the central charges (25a,25b) can be independently derived from the logarithmic divergent term of action by using the perturbation solution of order . One can consider general boundary conditions by adding intrinsic curvatures on Chu:2017aab (). In this case the boundary central charges change but the relations (15), (17) remain the same. We can also reproduce these relations in the holographic model Miao:2017gyt (); Chu:2017aab (). These are all strong supports for the universal relations (15), (17). The fact that the both the holographic models of Takayanagi:2011zk () and ours Miao:2017gyt (); Chu:2017aab () verify the universal relations (15), (17) suggests that both proposals are consistent holographic models of BCFT. We remark that in general there could be more than one self-consistent boundary conditions for a theory Song:2016pwx () and so there is no contradiction between Takayanagi:2011zk () and Miao:2017gyt (); Chu:2017aab (). This is supported by the fact that the two holographic models gives different boundary central charges despite the same universal relations are satisfied.

From holographic BCFT Takayanagi:2011zk (); Miao:2017gyt (); Chu:2017aab (), we can also gain some insight into the total energy. Applying the holographic renormalization of BCFT Miao:2017gyt (); Chu:2017aab (), we obtain the total stress tensor:

 Tij=2α1¯kijxd−1−δ(x;P)2α1d−2¯kijϵd−2+O(k2),x∼ϵ. (26)

Note that the first term, a local energy density, give rises to a divergence in the total energy that cannot be canceled with any local counterterm in the BCFT, but only with the inclusion of the second term, a surface counterterm as first constructed in Kennedy:1979ar (). The surface counterterm is localized at the boundary surface , which has been shifted from to a position . The requirement of finite energy fixes Kennedy:1979ar () the relative coefficients of the two terms in (26). Remarkably the holographic constructions Takayanagi:2011zk (); Miao:2017gyt (); Chu:2017aab () reproduce precisely also the surface counter term with the needed coefficient to make the total energy finite : , which agrees with the results of Kennedy:1979ar (); Dowker:1978md ().

## 4 Conclusions and Discussions

In this letter, we have shown that with the help of an effective action description, the divergent parts of the stress tensor of a BCFT is completely determined by the central charges of the theory. The found relations between the Casimir coefficients and the central charges are verified by free BCFT as well as holographic models of BCFT. We propose that these relations hold universally for any BCFT. Using the holographic models, we also reproduce remarkably the precise surface counterterm that is needed to render the total energy of the BCFT finite.

Our results are useful for the study of shape dependence of Casimir effects Emig:2001dx (); Schaden:2009zza (); Rajabpour:2016iwf () and the theory of BCFT Cardy:2004hm (); McAvity:1993ue (). For Casimir effects where there are spacetime on both sides of the boundary, it has been argued that the divergent stress tensor originates from the unphysical nature of classical “perfect conductor” boundary conditions Deutsch:1978sc (). In reality there would be an effective cut off below which the short wavelength vibrational modes do not “see the boundary”. However for BCFT where there is no spacetime outside the boundary, the divergent one point function of stress tensor is expected and physical. According to Cardy (), one can derive the one point function of an operator in BCFT from the two point functions of operators in CFT by using the mirror method. Since two point functions are divergent when two points are approaching, it is not surprising that the one point function of BCFT diverge near the boundary. This is due to the interaction with the boundary, or equivalently, the mirror image. Note that although the stress tensor diverges, the total energy is finite. Thus BCFT is self-consistent.

Our discussions can be generalized to higher dimensions naturally. Furthermore, our discussions also apply to defect conformal field theory (DCFT) Billo:2016cpy () with general codimensions, which is a problem of great interest. For example, the case of codimension 2 DCFT is related to the shape dependence of Rényi entropy Lewkowycz:2014jia (); Dong:2016wcf (); Chu:2016tps (); Bianchi:2016xvf (); Bianchi:2015liz (); Balakrishnan:2016ttg (). It is interesting to see whether the spirit of this letter can apply to general QFT. It is also very interesting to generalize and apply the techniques of the holographic models to study the expectation value of current in boundary systems, e.g. edge current of topological materials.

## Acknowledgements

We thank John Cardy, WuZhong Guo, Hugh Osborn and Douglas Smith for useful discussions and comments. This work is supported in part by NCTS and the grant MOST 105-2811-M-007-021 of the Ministry of Science and Technology of Taiwan.

## Appendix A Solutions to holographic BCFT

Here we give details about solutions to the Einstein equations and the boundary conditions (19) to the next order in derivative expansion of the boundary metric (i.e. in the case of a flat boundary metric ). Consider the following ansatz for ,

where the functions and are of order . We require that

 f(0)=1,X(0)=0,Qab(0)=qab (28)

so that the metric of BCFT takes the form (11) in Gauss normal coordinates.

### a.1 3d BCFT

Let us first study the case . The generalization to higher dimensions is straightforward. For simplicity, we further set , where are constants. Substituting (A) into the Einstein equations, and using (28) to fix the integral constants, we obtain (22) and

 f(s)=1−α1(s−g(s)) Q11(s)=18[4q1(s2+2)−α21(k1−k2)2(s2−3)g(s)2 −2α21(k1−k2)2log(s2+1)+s(5α21(k1−k2)2s+4α2) +s(2α1(−5k21+8k2k1+k22)−4s(k21−k2k1−k22+q2)) −2g(s)(α1k21(3α1s+s2−5)+2α2(s2+1)) −2α1g(s)(k22(3s(α1+s)+1)+2k1k2(4−3α1s))], Q22(s)=18[4q2(s2+2)−α21(k1−k2)2(s2−3)g(s)2 +s(5α21(k1−k2)2s−4α2)−2α21(k1−k2)2log(s2+1) +s(4s(k21+k2k1−k22−q1)−2α1(k21−4k2k1+7k22)) +2g(s)(2α2(s2+1)−α1k21(3α1s+s2−1)) +2α1g(s)(k22(−3α1s+s2+7)+2k1k2(3α1s+2s2−2))], X(s)=14[−α21(k1−k2)2s2log(s2+1)−2α1(k1−k2)2s +α1(k1−k2)2g(s)(α1(s2+1)g(s)+2s(s−α1)+2) +s(α21(k1−k2)2s−2s(k21+k2k1+k22−q1−q2))], (29)

where and . A continuous solution of the Einstein equations is obtained by first rewriting (A) as function of and and then analytic continutate to the region. In this way, we get smooth as (23). The solution is parametrized by two free parameters and .

Next we solve (19) for the embedding function of in the above background. We obtain, for , the results

 x=−sinh(ρ)z+kcosh2ρ2(d−1)z2+c3z3+O(k3) (30)

with given by

 c3=−sinhρ24[7k21+4k2k1+7k22−4(q1+q2) +α21(k1−k2)2((2+cosh(2ρ))log(coth2ρ)−1)]. (31)

The boundary conditions (19) also restrict solutions (23) and fix the integral constants to be

 α1=−1θ,  α2=−α12k2, (32)

where is the angle between and the bulk boundary . It should be mentioned that, following our method, the above is independently obtained in a recent paper Seminara:2017hhh (). The derivation of (30)-(32) is straightforward. For simplicity, let us first focus on the leading order term. From dimensional analysis, the embedding function of takes the form with a dimensionless constant. Substituting the metric (A) and the embedding function of into the conditions (19), we get two independent equations at order

 sech5(ρ)(−8c2+cosh(2ρ)+1)k=0, (α1cosh2(ρ)(4tan−1(tanhρ2)+π)+8c2)¯kab=0.

Solving the above equations, we obtain and as shown in (LABEL:3dQapp1), (32). Similarly, we obtain and from (19) at order . It is remarkable that the conditions (19) fix the bulk metric and embedding function of at the same time.

Substituting (23), (A),(A.1), (32) into (24), we obtain the holographic stress tensor

 Tij=diag{ α1(k1−k2)2x,α1(k1−k2)x2−3α1(k1−k2)22x, (33) α1(k2−k1)x2−3α1(k1−k2)22x}.

It is remarkable that all the dependence got cancelled away and the stress tensor (33) takes exactly the expected form (3) with coefficients as listed in Table 3. Recall that in (3) is actually a tensor defined at instead of the boundary . It can be obtained from parallel transport of the extrinsic curvature at , i.e., Deutsch:1978sc ().

Further generalization of the our above results is possible. Let us discuss briefly the case of non-constant metric and extrinsic curvature . In this case, will include non-diagonal parts generally. These non-diagonal parts obey (3b) trivially, since by definition (24) automatically satisfy the traceless and divergenceless conditions (2), which fixs the non-diagonal parts of stress tensor as (3b) completely.

Another generalization is to have more general boundary conditions of holographic BCFT by adding intrinsic curvatures on Chu:2017aab (). For example, we consider

 I=∫N√G(R−2Λ)+2∫Q√γ(K−T−λRQ), (34)

with the Neumann boundary condition

 Kαβ−(K−T−λRQ)γαβ−2λRQαβ=0. (35)

Substituting the solutions (23) into (35), we can solve the embedding function of as (30) but with different parameter and different integration constants

 α1 = 12λsechρ/(1−2λtanhρ)−θ, α2 = −α12k2. (36)

Here . From (24), we can derive the holographic stress tensor which takes exactly the expected form (3). It is remarkable that although the central charge changes, the relations (15) remain invariant for holographic BCFT with general boundary conditions. The above discussions can be generalized to higher dimensions easily. The 4d solutions can be used to confirm the universal relations (17).

### a.2 4d BCFT

Now Let us consider the case . For simplicity, we also set , where are constants. Substituting (A) into the Einstein equations, and using (28) to fix the integral constants, we obtain

 f(s)=1+2α1−α1(s2+2)√s2+1, (37)
 X(s)=16s2(2(q1+q2+q3)−3(k1k2+k1k3+k3k2))−13(k21+k22+k23−k1k2−k1k3−k2k3)g1(s), Q11(s)=k21g2(s)+k22g3(s)+k1k2g4(s)18(s2+1)3/2, +13√s2+1[q1(2s2+√s2+1+2)+q2(−s2+√s2+1−1)+3α2((√s2+1−2)s2+2(√s2+1−1))], Q22(s)=k22g2(s)+k21g3(s)+k2k1g4(s)18(s2+1)3/2 +13√s2+1[q2(2s2+√s2+1+2)+q1(−s2+√s2+1−1)+3α3((√s2+1−2)s2+2(√s2+1−1))], Q33(s)=(k21+k22)g5(s)+k1k2g6(s)18(s2+1)3/2 −(q1+q2)(s2−√s2+1+1)+3(α2+α3)((√s2+1−2)s2+2(√s2+1−1))3√s2+1,

where are defined by

 g1(s)=α1(α1(8√s2+1+s2(log(s2+1)−4)−8)−2s2+4√s2+1−4)+s2 g2(s)=12(s2+1)(−s2+√s2+1−1)+36α1(s2+1)(−s2+2√s2+1−2) −α21(−86(√s2+1−1)+s2(22s2−71√s2+1+108)+6(s2+1)3/2log(s2+1)) g3(s)=−6(s2+1)(−s2+√s2+1−1)+6α1(s2+1)(−s2+2√s2+1−2) +α21(14(√s2+1−1)+s2(2s2+11√s2+1−12)−6(s2+1)3/2log(s2+1)) g4(s)=−12(s2+1)(−s2+√s2+1−1)−30α1(s2+1)(−s2+2√s2+1−2) +α21(22s4−86(√s2+1−1)+s2(108−71√s2+1)+6(s2+1)3/2log(s2+1)) g5(s)=−6(s2+1)(−s2+√s2+1−1)−6α1(s2+1)((2√s2+1−3)s2+2(√s2+1−1)) +α21(44(√s2+1−1)+8s2(7√s2+1−9)−6(s2+1)3/2log(s2+1)+s4(15√s2+1−28)) g6(s)=3(s2+1)((3√s2+1−8)s2+8(√s2+1−1)) +12α1(s2+1)((√s2+1−3)s2+4(√s2+1−1)) +α21(4s2(√s2+1−6)+28(√s2+1−1)+6(s2+1)3/2log(s2+1)+s4(4−15√s