The basic starting idea in this work is simple. It is in the generalizations and implementations where difficulties arise.
Starting in Minkowski space in a fixed given Lorentzian frame with spatial origin, the electric dipole moment is calculated from an integral over the (localized) charge distribution. If there is a shift, , in the origin, the dipole transforms as
The first generalization is formal and somewhat artificial: shortly it will become quite natural. We introduce, in addition to the electric dipole moment, the magnetic dipole moment (also obtained by an integral over the current distribution) and write
By allowing the displacement to take complex values, , Eq. (6.37), can be generalized to
We emphasize that this is done in a fixed Lorentz frame and only the origin is moved. In different Lorentz frames there will be different complex centers of charge.
Later, directly from the general asymptotic Maxwell field itself (satisfying the Maxwell equations), we define the asymptotic complex dipole moment and give its transformation law, including transformations between Lorentz frames. This yields a unique complex center of charge independent of the Lorentz frame.
In this section, we describe how a complex center of charge for asymptotically vanishing Maxwell fields in flat spacetime can be found by using the shear-free NGCs, constructed from solutions of the homogeneous good-cut equation, to transform certain Maxwell field components to zero. Although this serves as a good example for our later methods in asymptotically flat spacetimes, the reader may wish to skip ahead to Section 4, where we go directly to gravitational fields in a setting of greater generality.
Our first set of applications of shear-free NGCs comes from Maxwell theory in Minkowski space. We review the general theory of the behavior of asymptotically-flat or vanishing Maxwell fields assuming throughout that there is a nonvanishing total charge, . As stated in Section 2, the Maxwell field is described in terms of its complex tetrad components, (). In a Bondi coordinate/tetrad system the asymptotic integration is relatively simple [50, 38] resulting in the radial behavior (the peeling theorem):
The formal integration procedure is to take as an arbitrary function of () (the free broadcasting data), then integrate the second, for , with a time-independent spin-weight function of integration and finally integrate the first, for . Using a slight modification of this, namely from the spherical harmonic expansion, we obtain,. For later use, the complex dipole is written as . Note that the is defined relative to a given Bondi system. This is the analogue of a given origin for the calculations of the dipole moments of Eq. (6.37).
Later in this section it will be shown that we can find a unique complex world line, , (the world line associated with a shear-free NGC), that is closely related to the . From this complex world line we can define the intrinsic complex dipole moment, .
However, we first discuss a particular Maxwell field, , where one of its eigenvectors is a tangent field to a shear-free NGC. This solution, referred to as the complex Liénard–Wiechert field is the direct generalization of the ordinary Liénard–Wiechert field. Though it is a real solution in Minkowski space, it can be thought of as arising from a complex world line in complex Minkowski space.
The parametric form of the general NGC was given earlier by Eq. (3.2),geodesic coordinates . Note that while these coordinates are not Bondi coordinates, though, in the limit, at , they are. The associated (geodesic) tetrad is given as a function of these geodesic coordinates, but with Minkowskian components by Eqs. (3.56). We restrict ourselves to the special case of the coordinates and tetrad associated with the from a shear-free NGC. Though we are dealing with a real shear-free twisting congruence, the congruence, as we saw, is generated by a complex analytic world line in the complexified Minkowski space, . The complex parameter, , must in the end be chosen so that the ‘’ of Eq. (3.19) is real. The Minkowski metric and the spin coefficients associated with this geodesic system can be calculated  in the frame. Unfortunately, it must be stated parametrically, since the explicitly appears via the and can not be directly eliminated. (An alternate choice of these geodesic coordinates is to use the instead of the . Unfortunately, this leads to an analytic flat metric on the complexified Minkowski space, where the real spacetime is hard to find.)
The use and insight given by this coordinate/tetrad system is illustrated by its application to a special class of Maxwell fields. We consider, as mentioned earlier, the Maxwell field where one of its principle null vectors, , (an eigenvector of the Maxwell tensor, ), is a tangent vector of a shear-free NGC. Thus, it depends on the choice of the complex world line and is therefore referred to as the complex Liénard–Wiechert field. (If the world line was real it would lead to the ordinary Liénard–Wiechert field.) We emphasize that though the source can formally be thought of as a charge moving on the complex world line, the Maxwell field is a real field on real Minkowski space. It will have a real (distributional) source at the caustics of the congruence. Physically, its behavior is very similar to real Liénard–Wiechert fields, the essential difference is that the electric dipole is now replaced by the combined electric and magnetic dipoles. The imaginary part of the world line determines the magnetic dipole moment.
In the spin-coefficient version of the Maxwell equations, using the geodesic tetrad, the choice of as the principle null vector ‘congruence’ is just the statement that
This allows a very simple exact integration of the remaining Maxwell components .
The present section, included as an illustration of the general ideas and constructions in this work, is rather technical and complicated and can be omitted without loss of continuity.
The complex Liénard–Wiechert fields (which we again emphasize are real Maxwell fields) are formally given by the (geodesic) tetrad components of the Maxwell tensor in the null geodesic coordinate system (), Eq. (3.56). As the detailed calculations are long  and take us too far afield, we only give an outline here. The integration of the radial Maxwell equations leads to the asymptotic behavior,
The remaining unknowns, , are determined by the last of the Maxwell equations,
These remaining equations depend only on , which, in turn, is determined by . In other words, the solution is driven by the complex line, . As they now stand, Eqs. (3.61) appear to be difficult to solve, partially due to the implicit description of the .
Actually they are easily solved when the independent variables are changed, via Eq. (3.15), from to the complex . They become, after a bit of work,
Though we now have the exact solution, unfortunately it is in complex coordinates where virtually every term depends on the complex variable , via . This is a severe impediment to a full description and understanding of the solution in the real Minkowski space.
In order to understand its asymptotic behavior and physical content, one must transform it, via Eqs. (3.62) – (3.67), back to a Bondi coordinate/tetrad system. This can only be done by approximations. After a lengthy calculation , we find the Bondi peeling behaviorsmall deviation from the straight line, , i.e., by
Again to first order, Eqs. (3.73), (3.74) and (3.75) yield
Reversing the issue, if we had instead started with an exact complex Liénard–Wiechert field but now given in a Bondi coordinate/tetrad system and performed on it the transformations, Eqs. (2.10) and (3.65) to the geodesic system, it would have resulted in
This example was intended to show how physical meaning could be attached to the complex world line associated with a shear-free NGC. In this case and later in the case of asymptotically-flat spacetimes, when the GCF is singled out by either the Maxwell field or the gravitational field, it will be referred to it as a UCF. For either of the two cases, a flat-space asymptotically-vanishing Maxwell field (with nonvanishing total charge) and for a vacuum asymptotically-flat spacetime, there will be a unique UCF. In the case of the Einstein–Maxwell fields there will, in general, be two UCFs: one for each field.
We return now to the general asymptotically-vanishing Maxwell field, Eqs. (3.48) and (3.51), and its transformation behavior under the null rotation around ,
The ‘picture’ to adopt is that the new s are now given in a tetrad defined by the complex light cone (or generalized light cone) with origin on the complex world line. (This is obviously formal and perhaps physically nonsensical, but mathematically quite sound, as the shear-free congruence can be thought of as having its origin on the complex line, .) From the physical identifications of charge, dipole moments, etc., of Eq. (3.54), we can obtain the transformation law of these physical quantities. In particular, the harmonic of , or, equivalently, the complex dipole, transforms asextract only the harmonic from a Clebsch–Gordon expansion of . A subtlety and difficulty of this extraction process is here clarified.
An important observation, obvious but easily overlooked, concerning the spherical harmonic expansions is that, in a certain sense, they lack uniqueness. As this issue is significant, its clarification is important.
Assume that we have a particular spin- function on , say, , given in a specific Bondi coordinate system, , that has a harmonic expansion given, for constant , by
If exactly the same function was given on different cuts or slices, say,
the harmonic expansion at constant would be different. The new coefficients are extracted by the two-sphere integral taken at constant :
The transformation, Eq. (3.84), and harmonic extraction implemented by first replacing the in all the terms of all , by , yields with a functional form ,
Though it is clear that extracting with this relationship is available in principle, in practice it is impossible to do it exactly and all examples are done with approximations: essentially using slow motion for the complex world line.
Remark 9. If by some accident the Maxwell field was a complex Liénard–Wiechert field, a world line could be chosen so that from the associated complex null cones we would have . However, though this cannot be done in general, the harmonics of can be made to vanish by the appropriate choice of the . This is the means by which a unique world line is chosen.
The complex center of charge is defined by the vanishing of the complex dipole moment ; in other words,
From Eq. (3.84),
Carrying this calculation  to second order, we find the second-order complex center of charge and the relationship between the intrinsic complex dipole, , and the complex dipole, ,
with this uniquely determined world line is referred to as the Maxwell UCF.
In Section 5, these ideas are applied to GR, with the complex electric and magnetic dipoles being replaced by the complex combination of the mass dipole and the angular momentum.
Living Rev. Relativity 15, (2012), 1
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