Let ${u: {\bf R}^3 \rightarrow {\bf R}^3}$ be a divergence-free vector field, thus ${\nabla \cdot u = 0}$, which we interpret as a velocity field. In this post we will proceed formally, largely ignoring the analytic issues of whether the fields in question have sufficient regularity and decay to justify the calculations. The vorticity field ${\omega: {\bf R}^3 \rightarrow {\bf R}^3}$ is then defined as the curl of the velocity:

$\displaystyle \omega = \nabla \times u.$

(From a differential geometry viewpoint, it would be more accurate (especially in other dimensions than three) to define the vorticity as the exterior derivative ${\omega = d(g \cdot u)}$ of the musical isomorphism ${g \cdot u}$ of the Euclidean metric ${g}$ applied to the velocity field ${u}$; see these previous lecture notes. However, we will not need this geometric formalism in this post.)

Assuming suitable regularity and decay hypotheses of the velocity field ${u}$, it is possible to recover the velocity from the vorticity as follows. From the general vector identity ${\nabla \times \nabla \times X = \nabla(\nabla \cdot X) - \Delta X}$ applied to the velocity field ${u}$, we see that

$\displaystyle \nabla \times \omega = -\Delta u$

and thus (by the commutativity of all the differential operators involved)

$\displaystyle u = - \nabla \times \Delta^{-1} \omega.$

Using the Newton potential formula

$\displaystyle -\Delta^{-1} \omega(x) := \frac{1}{4\pi} \int_{{\bf R}^3} \frac{\omega(y)}{|x-y|}\ dy$

and formally differentiating under the integral sign, we obtain the Biot-Savart law

$\displaystyle u(x) = \frac{1}{4\pi} \int_{{\bf R}^3} \frac{\omega(y) \times (x-y)}{|x-y|^3}\ dy. \ \ \ \ \ (1)$

This law is of fundamental importance in the study of incompressible fluid equations, such as the Euler equations

$\displaystyle \partial_t u + (u \cdot \nabla) u = -\nabla p; \quad \nabla \cdot u = 0$

since on applying the curl operator one obtains the vorticity equation

$\displaystyle \partial_t \omega + (u \cdot \nabla) \omega = (\omega \cdot \nabla) u \ \ \ \ \ (2)$

and then by substituting (1) one gets an autonomous equation for the vorticity field ${\omega}$. Unfortunately, this equation is non-local, due to the integration present in (1).

In a recent work, it was observed by Elgindi that in a certain regime, the Biot-Savart law can be approximated by a more “low rank” law, which makes the non-local effects significantly simpler in nature. This simplification was carried out in spherical coordinates, and hinged on a study of the invertibility properties of a certain second order linear differential operator in the latitude variable ${\theta}$; however in this post I would like to observe that the approximation can also be seen directly in Cartesian coordinates from the classical Biot-Savart law (1). As a consequence one can also initiate the beginning of Elgindi’s analysis in constructing somewhat regular solutions to the Euler equations that exhibit self-similar blowup in finite time, though I have not attempted to execute the entirety of the analysis in this setting.

Elgindi’s approximation applies under the following hypotheses:

A model example of a divergence-free vector field obeying these properties (but without good decay at infinity) is the linear vector field

$\displaystyle X(x) = (x_1, x_2, -2x_3) \ \ \ \ \ (5)$

which is of the form (3) with ${u_r(r,x_3) = r}$ and ${u_3(r,x_3) = -2x_3}$. The associated vorticity ${\omega}$ vanishes.

We can now give an illustration of Elgindi’s approximation:

Proposition 1 (Elgindi’s approximation) Under the above hypotheses (and assuing suitable regularity and decay), we have the pointwise bounds

$\displaystyle u(x) = \frac{1}{2} {\mathcal L}_{12}(\omega)(|x|) X(x) + O( |x| \|\omega\|_{L^\infty({\bf R}^3)} )$

for any ${x \in {\bf R}^3}$, where ${X}$ is the vector field (5), and ${{\mathcal L}_{12}(\omega): {\bf R}^+ \rightarrow {\bf R}}$ is the scalar function

$\displaystyle {\mathcal L}_{12}(\omega)(\rho) := \frac{3}{4\pi} \int_{|y| \geq \rho} \frac{r y_3}{|y|^5} \omega_{r3}(r,y_3)\ dy.$

Thus under the hypotheses (i), (ii), and assuming that ${\omega}$ is slowly varying, we expect ${u}$ to behave like the linear vector field ${X}$ modulated by a radial scalar function. In applications one needs to control the error in various function spaces instead of pointwise, and with ${\omega}$ similarly controlled in other function space norms than the ${L^\infty}$ norm, but this proposition already gives a flavour of the approximation. If one uses spherical coordinates

$\displaystyle \omega_{r3}( \rho \cos \theta, \rho \sin \theta ) = \Omega( \rho, \theta )$

then we have (using the spherical change of variables formula ${dy = \rho^2 \cos \theta d\rho d\theta d\phi}$ and the odd nature of ${\Omega}$)

$\displaystyle {\mathcal L}_{12}(\omega) = L_{12}(\Omega),$

where

$\displaystyle L_{12}(\Omega)(\rho) = 3 \int_\rho^\infty \int_0^{\pi/2} \frac{\Omega(r, \theta) \sin(\theta) \cos^2(\theta)}{r}\ d\theta dr$

is the operator introduced in Elgindi’s paper.

Proof: By a limiting argument we may assume that ${x}$ is non-zero, and we may normalise ${\|\omega\|_{L^\infty({\bf R}^3)}=1}$. From the triangle inequality we have

$\displaystyle \int_{|y| \leq 10|x|} \frac{\omega(y) \times (x-y)}{|x-y|^3}\ dy \leq \int_{|y| \leq 10|x|} \frac{1}{|x-y|^2}\ dy$

$\displaystyle \leq \int_{|z| \leq 11 |x|} \frac{1}{|z|^2}\ dz$

$\displaystyle = O( |x| )$

and hence by (1)

$\displaystyle u(x) = \frac{1}{4\pi} \int_{|y| > 10|x|} \frac{\omega(y) \times (x-y)}{|x-y|^3}\ dy + O(|x|).$

In the regime ${|y| > 2|x|}$ we may perform the Taylor expansion

$\displaystyle \frac{x-y}{|x-y|^3} = \frac{x-y}{|y|^3} (1 - \frac{2 x \cdot y}{|y|^2} + \frac{|x|^2}{|y|^2})^{-3/2}$

$\displaystyle = \frac{x-y}{|y|^3} (1 + \frac{3 x \cdot y}{|y|^2} + O( \frac{|x|^2}{|y|^2} ) )$

$\displaystyle = -\frac{y}{|y|^3} + \frac{x}{|y|^3} - \frac{3 (x \cdot y) y}{|y|^5} + O( \frac{|x|^2}{|y|^4} ).$

Since

$\displaystyle \int_{|y| > 10|x|} \frac{|x|^2}{|y|^4}\ dy = O(|x|)$

we see from the triangle inequality that the error term contributes ${O(|x|)}$ to ${u(x)}$. We thus have

$\displaystyle u(x) = -A_0(x) + A_1(x) - 3A'_1(x) + O(|x|)$

where ${A_0}$ is the constant term

$\displaystyle A_0 := \int_{|y| > 10|x|} \frac{\omega(y) \times y}{|y|^3}\ dy,$

and ${A_1, A'_1}$ are the linear term

$\displaystyle A_1 := \int_{|y| > 10|x|} \frac{\omega(y) \times x}{|y|^3}\ dy,$

$\displaystyle A'_1 := \int_{|y| > 10|x|} (x \cdot y) \frac{\omega(y) \times y}{|y|^5}\ dy.$

By the hypotheses (i), (ii), we have the symmetries

$\displaystyle \omega(y_1,y_2,-y_3) = - \omega(y_1,y_2,y_3) \ \ \ \ \ (6)$

and

$\displaystyle \omega(-y_1,-y_2,y_3) = - \omega(y_1,y_2,y_3) \ \ \ \ \ (7)$

and hence also

$\displaystyle \omega(-y_1,-y_2,-y_3) = \omega(y_1,y_2,y_3). \ \ \ \ \ (8)$

The even symmetry (8) ensures that the integrand in ${A_0}$ is odd, so ${A_0}$ vanishes. The symmetry (6) or (7) similarly ensures that ${\int_{|y| > 10|x|} \frac{\omega(y)}{|y|^3}\ dy = 0}$, so ${A_1}$ vanishes. Since ${\int_{|x| < y \leq 10|x|} \frac{|x \cdot y| |y|}{|y|^5}\ dy = O( |x| )}$, we conclude that

$\displaystyle \omega(x) = -3\int_{|y| \geq |x|} (x \cdot y) \frac{\omega(y) \times y}{|y|^5}\ dy + O(|x|).$

Using (4), the right-hand side is

$\displaystyle -3\int_{|y| \geq |x|} (x_1 y_1 + x_2 y_2 + x_3 y_3) \frac{\omega_{r3}(r,y_3) (-y_1 y_3, -y_2 y_3, y_1^2+y_2^2)}{r|y|^5}\ dy$

$\displaystyle + O(|x|)$

where ${r := \sqrt{y_1^2+y_2^2}}$. Because of the odd nature of ${\omega_{r3}}$, only those terms with one factor of ${y_3}$ give a non-vanishing contribution to the integral. Using the rotation symmetry ${(y_1,y_2,y_3) \mapsto (-y_2,y_1,y_3)}$ we also see that any term with a factor of ${y_1 y_2}$ also vanishes. We can thus simplify the above expression as

$\displaystyle -3\int_{|y| \geq |x|} \frac{\omega_{r3}(r,y_3) (-x_1 y_1^2 y_3, -x_2 y_2^2 y_3, x_3 (y_1^2+y_2^2) y_3)}{r|y|^5}\ dy + O(|x|).$

Using the rotation symmetry ${(y_1,y_2,y_3) \mapsto (-y_2,y_1,y_3)}$ again, we see that the term ${y_1^2}$ in the first component can be replaced by ${y_2^2}$ or by ${\frac{1}{2} (y_1^2+y_2^2) = \frac{r^2}{2}}$, and similarly for the ${y_2^2}$ term in the second component. Thus the above expression is

$\displaystyle \frac{3}{2} \int_{|y| \geq |x|} \frac{\omega_{r3}(r,y_3) (x_1 , x_2, -2x_3) r y_3}{|y|^5}\ dy + O(|x|)$

giving the claim. $\Box$

Example 2 Consider the divergence-free vector field ${u := \nabla \times \psi}$, where the vector potential ${\psi}$ takes the form

$\displaystyle \psi(x_1,x_2,x_3) := (x_2 x_3, -x_1 x_3, 0) \eta(|x|)$

for some bump function ${\eta: {\bf R} \rightarrow {\bf R}}$ supported in ${(0,+\infty)}$. We can then calculate

$\displaystyle u(x_1,x_2,x_3) = X(x) \eta(|x|) + (x_1 x_3, x_2 x_3, -x_1^2-x_2^2) \frac{\eta'(|x|) x_3}{|x|}.$

and

$\displaystyle \omega(x_1,x_2,x_3) = (-6x_2 x_3, 6x_1 x_3, 0) \frac{\eta'(|x|)}{|x|} + (-x_2 x_3, x_1 x_3, 0) \eta''(|x|).$

In particular the hypotheses (i), (ii) are satisfied with

$\displaystyle \omega_{r3}(r,x_3) = - 6 \eta'(|x|) \frac{x_3 r}{|x|} - \eta''(|x|) x_3 r.$

One can then calculate

$\displaystyle L_{12}(\omega)(\rho) = -\frac{3}{4\pi} \int_{|y| \geq \rho} (6\frac{\eta'(|y|)}{|y|^6} + \frac{\eta''(|y|)}{|y|^5}) r^2 y_3^2\ dy$

$\displaystyle = -\frac{2}{5} \int_\rho^\infty 6\eta'(s) + s\eta''(s)\ ds$

$\displaystyle = 2\eta(\rho) + \frac{2}{5} \rho \eta'(\rho).$

If we take the specific choice

$\displaystyle \eta(\rho) = \varphi( \rho^\alpha )$

where ${\varphi}$ is a fixed bump function supported some interval ${[c,C] \subset (0,+\infty)}$ and ${\alpha>0}$ is a small parameter (so that ${\eta}$ is spread out over the range ${\rho \in [c^{1/\alpha},C^{1/\alpha}]}$), then we see that

$\displaystyle \| \omega \|_{L^\infty} = O( \alpha )$

(with implied constants allowed to depend on ${\varphi}$),

$\displaystyle L_{12}(\omega)(\rho) = 2\eta(\rho) + O(\alpha),$

and

$\displaystyle u = X(x) \eta(|x|) + O( \alpha |x| ),$

which is completely consistent with Proposition 1.

One can use this approximation to extract a plausible ansatz for a self-similar blowup to the Euler equations. We let ${\alpha>0}$ be a small parameter and let ${\omega_{rx_3}}$ be a time-dependent vorticity field obeying (i), (ii) of the form

$\displaystyle \omega_{rx_3}(t,r,x_3) \approx \alpha \Omega( t, R ) \mathrm{sgn}(x_3)$

where ${R := |x|^\alpha = (r^2+x_3^2)^{\alpha/2}}$ and ${\Omega: {\bf R} \times [0,+\infty) \rightarrow {\bf R}}$ is a smooth field to be chosen later. Admittedly the signum function ${\mathrm{sgn}}$ is not smooth at ${x_3}$, but let us ignore this issue for now (to rigorously make an ansatz one will have to smooth out this function a little bit; Elgindi uses the choice ${(|\sin \theta| \cos^2 \theta)^{\alpha/3} \mathrm{sgn}(x_3)}$, where ${\theta := \mathrm{arctan}(x_3/r)}$). With this ansatz one may compute

$\displaystyle {\mathcal L}_{12}(\omega(t))(\rho) \approx \frac{3\alpha}{2\pi} \int_{|y| \geq \rho; y_3 \geq 0} \Omega(t,R) \frac{r y_3}{|y|^5}\ dy$

$\displaystyle = \alpha \int_\rho^\infty \Omega(t, s^\alpha) \frac{ds}{s}$

$\displaystyle = \int_{\rho^\alpha}^\infty \Omega(t,s) \frac{ds}{s}.$

By Proposition 1, we thus expect to have the approximation

$\displaystyle u(t,x) \approx \frac{1}{2} \int_{|x|^\alpha}^\infty \Omega(t,s) \frac{ds}{s} X(x).$

We insert this into the vorticity equation (2). The transport term ${(u \cdot \nabla) \omega}$ will be expected to be negligible because ${R}$, and hence ${\omega_{rx_3}}$, is slowly varying (the discontinuity of ${\mathrm{sgn}(x_3)}$ will not be encountered because the vector field ${X}$ is parallel to this singularity). The modulating function ${\frac{1}{2} \int_{|x|^\alpha}^\infty \Omega(t,s) \frac{ds}{s}}$ is similarly slowly varying, so derivatives falling on this function should be lower order. Neglecting such terms, we arrive at the approximation

$\displaystyle (\omega \cdot \nabla) u \approx \frac{1}{2} \int_{|x|^\alpha}^\infty \Omega(t,s) \frac{ds}{s} \omega$

and so in the limit ${\alpha \rightarrow 0}$ we expect obtain a simple model equation for the evolution of the vorticity envelope ${\Omega}$:

$\displaystyle \partial_t \Omega(t,R) = \frac{1}{2} \int_R^\infty \Omega(t,S) \frac{dS}{S} \Omega(t,R).$

If we write ${L(t,R) := \int_R^\infty \Omega(t,S)\frac{dS}{S}}$ for the logarithmic primitive of ${\Omega}$, then we have ${\Omega = - R \partial_R L}$ and hence

$\displaystyle \partial_t (R \partial_R L) = \frac{1}{2} L (R \partial_R L)$

which integrates to the Ricatti equation

$\displaystyle \partial_t L = \frac{1}{4} L^2$

which can be explicitly solved as

$\displaystyle L(t,R) = \frac{2}{f(R) - t/2}$

where ${f(R)}$ is any function of ${R}$ that one pleases. (In Elgindi’s work a time dilation is used to remove the unsightly factor of ${1/2}$ appearing here in the denominator.) If for instance we set ${f(R) = 1+R}$, we obtain the self-similar solution

$\displaystyle L(t,R) = \frac{2}{1+R-t/2}$

and then on applying ${-R \partial_R}$

$\displaystyle \Omega(t,R) = \frac{2R}{(1+R-t/2)^2}.$

Thus, we expect to be able to construct a self-similar blowup to the Euler equations with a vorticity field approximately behaving like

$\displaystyle \omega(t,x) \approx \alpha \frac{2R}{(1+R-t/2)^2} \mathrm{sgn}(x_3) (\frac{x_2}{r}, -\frac{x_1}{r}, 0)$

and velocity field behaving like

$\displaystyle u(t,x) \approx \frac{1}{1+R-t/2} X(x).$

In particular, ${u}$ would be expected to be of regularity ${C^{1,\alpha}}$ (and smooth away from the origin), and blows up in (say) ${L^\infty}$ norm at time ${t/2 = 1}$, and one has the self-similarity

$\displaystyle u(t,x) = (1-t/2)^{\frac{1}{\alpha}-1} u( 0, \frac{x}{(1-t/2)^{1/\alpha}} )$

and

$\displaystyle \omega(t,x) = (1-t/2)^{-1} \omega( 0, \frac{x}{(1-t/2)^{1/\alpha}} ).$

A self-similar solution of this approximate shape is in fact constructed rigorously in Elgindi’s paper (using spherical coordinates instead of the Cartesian approach adopted here), using a nonlinear stability analysis of the above ansatz. It seems plausible that one could also carry out this stability analysis using this Cartesian coordinate approach, although I have not tried to do this in detail.