In this section we consider the motion of a point particle of mass subjected to its own gravitational field. The particle moves on a world line in a curved spacetime whose background metric is assumed to be a vacuum solution to the Einstein field equations. We shall suppose that is small, so that the perturbation created by the particle can also be considered to be small; it will obey a linear wave equation in the background spacetime. This linearization of the field equations will allow us to fit the problem of determining the motion of a point mass within the framework developed in Sections 5.1 and 5.2, and we shall obtain the equations of motion by following the same general line of reasoning. We shall find that is not a geodesic of the background spacetime because acts on the particle and induces an acceleration of order ; the motion is geodesic in the test-mass limit only.
Our discussion in this first section is largely formal: As in Sections 5.1.1 and 5.2.1 we insert the point particle in the background spacetime and ignore the fact that the field it produces is singular on the world line. To make sense of the formal equations of motion will be our goal in the following Sections 5.3.2, 5.3.3, 5.3.4, 5.3.5, 5.3.6, and 5.3.7. The problem of determining the motion of a small mass in a background spacetime will be reconsidered in Section 5.4 from a different and more satisfying premise: There the small body will be modeled as a black hole instead of as a point particle, and the singular behaviour of the perturbation will automatically be eliminated.
Let a point particle of mass move on a world line in a curved spacetime with metric . This is the total metric of the perturbed spacetime, and it depends on as well as all other relevant parameters. At a later stage of the discussion the total metric will be broken down into a “background” part that is independent of , and a “perturbation” part that is proportional to . The world line is described by relations in which is an arbitrary parameter – this will later be identified with proper time in the background spacetime. In this and the following sections we will use symbols to denote tensors that refer to the perturbed spacetime; tensors in the background spacetime will be denoted, as usual, by italic symbols.
The particle’s action functional isonly source of matter in the spacetime – an explanation will be provided at the end of this section – so that the Einstein field equations take the form of
On a formal level the metric is obtained by solving the Einstein field equations, and the world line is determined by solving the equations of energy-momentum conservation, which follow from the field equations. From Equations (81, 260, 484) we obtain
and additional manipulations reduce this to
where is the covariant acceleration and is a scalar field on the world line. Energy-momentum conservation therefore produces the geodesic equation
At this stage we begin treating as a formal expansion parameter, and we write
We have already stated that the particle is the only source of matter in the spacetime, and the metric must therefore be a solution to the vacuum field equations: . Equations (483, 488, (489) then imply , in which both sides of the equation are of order . To simplify the expression of the first-order correction to the Einstein tensor we introduce the trace-reversed gravitational potentials
The equations of motion for the point mass are obtained by substituting the expansion of Equation (487) into Equations (485) and (486). The perturbed connection is easily computed to be , and this leads to
having once more selected proper time (as measured in the background spacetime) as the parameter on the world line. On the other hand, Equation (486) gives
where is the particle’s acceleration vector. Since it is clear that the acceleration will be of order , the second term can be discarded and we obtain
Keeping the error term implicit, we shall express this in the equivalent form
It should be clear that Equation (494) is valid only in a formal sense, because the potentials obtained from Equations (493) diverge on the world line. The nonlinearity of the Einstein field equations makes this problem even worse here than for the scalar and electromagnetic cases, because the singular behaviour of the perturbation might render meaningless a formal expansion of in powers of . Ignoring this issue for the time being (we shall return to it in Section 5.4), we will proceed as in Sections 5.1 and 5.2 and attempt, with a careful analysis of the field’s singularity structure, to make sense of these equations.
To conclude this section I should explain why it is desirable to restrict our discussion to spacetimes that contain no matter except for the point particle. Suppose, in contradiction with this assumption, that the background spacetime contains a distribution of matter around which the particle is moving. (The corresponding vacuum situation has the particle moving around a black hole. Notice that we are still assuming that the particle moves in a region of spacetime in which there is no matter; the issue is whether we can allow for a distribution of matter somewhere else.) Suppose also that the matter distribution is described by a collection of matter fields . Then the field equations satisfied by the matter have the schematic form , and the metric is determined by the Einstein field equations , in which stands for the matter’s stress-energy tensor. We now insert the point particle in the spacetime, and recognize that this displaces the background solution to a new solution (. The perturbations are determined by the coupled set of equations and . After linearization these take the form of
where , , , and are suitable differential operators acting on the perturbations. This is a coupled set of partial differential equations for the perturbations and . These equations are linear, but they are much more difficult to deal with than the single equation for that was obtained in the vacuum case. And although it is still possible to solve the coupled set of equations via a Green’s function technique, the degree of difficulty is such that we will not attempt this here. We shall, therefore, continue to restrict our attention to the case of a point particle moving in a vacuum (globally Ricci-flat) background spacetime.
The retarded solution to Equation (493) is , where is the retarded Green’s function introduced in Section 4.5. After substitution of the stress-energy tensor of Equation (490) we obtain
For a more concrete expression we must take to be in a neighbourhood of the world line. The following manipulations follow closely those performed in Section 5.1.2 for the case of a scalar charge, and in Section 5.2.2 for the case of an electric charge. Because these manipulations are by now familiar, it will be sufficient here to present only the main steps. There are two important simplifications that occur in the case of a massive particle. First, for the purposes of computing to first order in , it is sufficient to take the world line to be a geodesic of the background spacetime: The deviations from geodesic motion that we are in the process of calculating are themselves of order and would affect at order only. We shall therefore be allowed to set
With the understanding that is close to the world line (refer back to Figure 9), we substitute the Hadamard construction of Equation (352) into Equation (495) and integrate over the portion of that is contained in . The result is
In the following Sections 5.3.3, 5.3.4, 5.3.5, 5.3.6, and 5.3.7, we shall refer to as the gravitational potentials at produced by a particle of mass moving on the world line , and to as the gravitational field at . To compute this is our next task.
Keeping in mind that and are related by , a straightforward computation reveals that the covariant derivatives of the gravitational potentials are given by
We wish to express in the retarded coordinates of Section 3.3, as an expansion in powers of . For this purpose we decompose the field in the tetrad that is obtained by parallel transport of on the null geodesic that links to ; this construction is detailed in Section 3.3. Note that throughout this section we set , where is the rotation tensor defined by Equation (138): The tetrad vectors are taken to be parallel transported on . We recall from Equation (141) that the parallel propagator can be expressed as . The expansion relies on Equation (166) for and Equation (168) for , both specialized to the case of geodesic motio, . We shall also need
Making these substitutions in Equation (484) and projecting against various members of the tetrad gives
The translation of the results contained in Equations (505, 506, 507, 508, 509, 510) into the Fermi normal coordinates of Section 3.2 proceeds as in Sections 5.1.4 and 5.2.4, but is simplified by the fact that here the world line can be taken to be a geodesic. We may thus set in Equations (224) and (225) that relate the tetrad to , as well as in Equations (221, 222, 223) that relate the Fermi normal coordinates to the retarded coordinates. We recall that the Fermi normal coordinates refer to a point on the world line that is linked to by a spacelike geodesic that intersects orthogonally.
The translated results are
It is then a simple matter to average these results over a two-surface of constant and . Using the area element of Equation (404) and definitions analogous to those of Equation (405), we obtain
The singular gravitational potentials
To evaluate the integral of Equation (525) we take to be close to the world line (see Figure 9), and we invoke Equation (373) as well as the Hadamard construction of Equation (379). This gives
Differentiation of Equation (526) yields
To derive an expansion for we follow the general method of Section 3.4.4 and introduce the functions . We have that
where overdots indicate differentiation with respect to and . The leading term was worked out in Equation (501), and the derivatives of are given by
according to Equations (503) and (360). Combining these results together with Equation (229) for gives
We proceed similarly to obtain an expansion for . Here we introduce the functions and express as . The leading term was computed in Equation (502), and
follows from Equation (359). Combining these results together with Equation (229) for gives
We obtain the frame components of the singular gravitational field by substituting these expansions into Equation (527) and projecting against the tetrad . After some algebra we arrive at
The difference between the retarded field of Equations (505, 506, 507, 508, 509, 510) and the singular field of Equations (532, 533, 534, 535, 536, 537) defines the radiative gravitational field . Its tetrad components are
The retarded gravitational field of a point particle is singular on the world line, and this behaviour makes it difficult to understand how the field is supposed to act on the particle and influence its motion. The field’s singularity structure was analyzed in Sections 5.3.3 and 5.3.4, and in Section 5.3.5 it was shown to originate from the singular field ; the radiative field was then shown to be smooth on the world line.
To make sense of the retarded field’s action on the particle we can follow the discussions of Section 5.1.6 and 5.2.6 and postulate that the self gravitational field of the point particle is either , as worked out in Equation (523), or , as worked out in Equation (544). These regularized fields are both given by
The actual gravitational perturbation is obtained by inverting Equation (491), which leads to . Substituting Equation (546) yields
Equation (550) was first derived by Yasushi Mino, Misao Sasaki, and Takahiro Tanaka in 1997 . (An incomplete treatment had been given previously by Morette-DeWitt and Ging .) An alternative derivation was then produced, also in 1997, by Theodore C. Quinn and Robert M. Wald . These equations are now known as the MiSaTaQuWa equations of motion. It should be noted that Equation (550) is formally equivalent to the statement that the point particle moves on a geodesic in a spacetime with metric , where is the radiative metric perturbation obtained by trace-reversal of the potentials ; this perturbed metric is smooth on the world line, and it is a solution to the vacuum field equations. This elegant interpretation of the MiSaTaQuWa equations was proposed in 2002 by Steven Detweiler and Bernard F. Whiting . Quinn and Wald  have shown that under some conditions, the total work done by the gravitational self-force is equal to the energy radiated (in gravitational waves) by the particle.
The equations of motion derived in the preceding Section 5.3.6 refer to a specific choice of gauge for the metric perturbation produced by a point particle of mass . We indeed recall that back at Equation (492) we imposed the Lorenz gauge condition on the gravitational potentials . By virtue of this condition we found that the potentials satisfy the wave equation of Equation (493) in a background spacetime with metric . The hyperbolic nature of this equation allowed us to identify the retarded solution as the physically relevant solution, and the equations of motion were obtained by removing the singular part of the retarded field. It seems clear that the Lorenz condition is a most appropriate choice of gauge.
Once the equations of motion have been formulated, however, the freedom of performing a gauge transformation (either away from the Lorenz gauge, or within the class of Lorenz gauges) should be explored. A gauge transformation will affect the form of the equations of motion: These must depend on the choice of coordinates, and there is no reason to expect Equation (550) to be invariant under a gauge transformation. Our purpose in this section is to work out how the equations of motion change under such a transformation. This issue was first examined by Barack and Ori .
We introduce a coordinate transformation of the form
and this change can be interpreted as a gauge transformation of the metric perturbation created by the moving particle:
To compute the gauge acceleration we substitute Equation (552) into Equation (494), and we simplify the result by invoking Ricci’s identity, , and the fact that . The final expression is
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