Suppose we take our buffer-region expansion of to be valid everywhere in the interior of (in ), rather than just in the buffer region. This is a meaningful supposition in a distributional sense, since the singularity in is locally integrable even at . Note that the extension of the buffer-region expansion is not intended to provide an accurate or meaningful approximation in the interior; it is used only as a means of determining the field in the exterior. We can do this because the field values in are entirely determined by the field values on , so using the buffer-region expansion in the interior of leaves the field values in unaltered. Now, given the extension of the buffer-region expansion, it follows from Stokes’ law that the integral over in Eq. (21.8) can be replaced by a volume integral over the interior of the tube, plus two surface integrals over the “caps” and , which fill the “holes” in and , respectively, where they intersect . Schematically, we can write Stokes’ law as , where is the interior of . This is now valid as a distributional identity. (Note that the “interior” here means the region bounded by ; does not refer to the set of interior points in the point-set defined by .) The minus sign in front of the integral over accounts for the fact that the directed surface element in Eq. (21.8) points into the tube. Because does not lie in the past of any point in , it does not contribute to the perturbation at . Hence, we can rewrite Eq. (21.8) as

where is the contribution from the spatial surface , and denotes the action of the wave-operator on . Now note that ; since , this term integrates to zero. Next note that vanishes everywhere except at . This means that the field at can be written as Making use of the fact that , along with the identity , where is a coordinate delta function in Fermi coordinates, we arrive at the desired result Therefore, in the region , the leading-order perturbation produced by the asymptotically small body is identical to the field produced by a point particle. At second order, the same method can be used to simplify Eq. (21.5) by replacing at least part of the integral over with an integral over . We will not pursue this simplification here, however.Gralla and Wald [83] provided an alternative derivation of the same result, using distributional methods to prove that the distributional source for the linearized Einstein equation must be that of a point particle in order for the solution to diverge as . One can understand this by considering that the most divergent term in the linearized Einstein tensor is a Laplacian acting on the perturbation, and the Laplacian of is a flat-space delta function; the less divergent corrections are due to the curvature of the background, which distorts the flat-space distribution into a covariant curved-spacetime distribution.

We have just seen that the solution to the wave equation with a point-mass source is given by

One can also obtain this result from Eq. (19.29) by taking the trace-reversal and making use of the Green’s function identity (16.22). In this section, we seek an expansion of the perturbation in Fermi coordinates. Following the same steps as in Section 19.2, we arrive at Here, primed indices now refer to the retarded point on the world line, is the retarded radial coordinate at , and the tail integral is given by where is the first time at which the world line enters , and denotes the time when it crosses the initial-data surface .We expand the direct term in in powers of using the following: the near-coincidence expansion ; the relationship between and , given by Eq. (11.5); and the coordinate expansion of the parallel-propagators, obtained from the formula , where the retarded tetrad can be expanded in terms of using Eqs. (11.9), (11.10), (9.12), and (9.13). We expand the tail integral similarly: noting that , we expand about as ; each term is then expanded using the near-coincidence expansions and , where barred indices correspond to the point , and is given by

This yields the expansion As with the direct part, the final coordinate expansion is found by expressing in terms of the Fermi tetrad.Combining the expansions of the direct and tail parts of the perturbation, we arrive at the expansion in Fermi coordinates:

As the final step, each of these terms is decomposed into irreducible STF pieces using the formulas (B.1), (B.3), and (B.7), yielding where the uppercase hatted tensors are specified in Table 2. Because the STF decomposition is unique, these tensors must be identical to the free functions in Eq. (22.38); hence, those free functions, comprising a regular, homogenous solution in the buffer region, have been uniquely determined by boundary conditions and waves emitted by the particle in the past.The Detweiler–Whiting singular field is given by

Using the Hadamard decomposition , we can write this as where primed indices now refer to the retarded point ; double-primed indices refer to the advanced point ; barred indices refer to points in the segment of the world line between and . The first term in Eq. (23.16) can be read off from the calculation of the retarded field. The other terms are expanded using the identities and . The final result isThe regular field could be calculated from the regular Green’s function. But it is more straightforwardly calculated using . The result is

With the metric perturbation fully determined, we can now express the self-force in terms of tail integrals. Reading off the components of from Table 2 and inserting the results into Eq. (22.84), we arrive at

We have now firmly established the results of the point-particle analysis.

Living Rev. Relativity 14, (2011), 7
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