3.1. Defining an umbilic surface and edge

The initial shape of a stellarator symmetric umbilic-torus-like surface $\partial V$ is parametrised as a rotating polygon-like with convex sides:

(3.1) \begin{align} R_{\mathrm{US}} & = R_0 + 2\varrho \cos \left (\frac {\pi }{2n}\right ) \cos \bigg(\frac {\theta + \frac {\pi }{n} (2 \lfloor n\theta/2\pi \rfloor + 1)}{2}\nonumber \\&\qquad\qquad\qquad\qquad\qquad + \frac {m n_{\textrm {FP}} \zeta + r_1 \sin (n_{\textrm {FP}} \zeta ) + r_4 \sin (2 n_{\textrm {FP}} \zeta )}{n} \bigg)\nonumber \\ & \quad - \varrho \cos \left (\frac {\pi }{n}\left(2 \Big \lfloor \frac {n \theta }{2\pi } \Big \rfloor + 1\right) + \frac {m n_{\textrm {FP}} \zeta + r_1 \sin (n_{\textrm {FP}} \zeta ) + r_4 \sin (2 n_{\textrm {FP}} \zeta )}{n}\right )\nonumber \\&\quad + r_2 \cos (n_{\textrm {FP}} \zeta ), \end{align}

(3.2) \begin{align} Z_{\mathrm{US}} &= 2\varrho \cos \left (\frac {\pi }{2n}\right ) \sin \left (\frac {\theta + \frac {\pi }{n} (2 \lfloor {n \theta }/{2\pi } \rfloor + 1)}{2} + \frac {m n_{\textrm {FP}} \zeta + r_5 \sin (n_{\textrm {FP}} \zeta )}{n} \right )\nonumber \\ & - \varrho \sin \left (\frac {\pi }{n}\left(2 \Big \lfloor \frac {n \theta }{2\pi } \Big \rfloor + 1\right) + \frac {m n_{\textrm {FP}} \zeta + r_5 \sin (n_{\textrm {FP}} \zeta )}{n}\right ) + r_3 \sin ( n_{\textrm {FP}} \zeta ), \end{align}

where the major radius $ R_0 = 1$ throughout this paper, unless stated otherwise, $\varrho$ is the minor radius of the umbilic-torus-like surface, $\theta$ is the poloidal angle used in DESC, $\zeta$ is the cylindrical toroidal angle and $\lfloor \ldots \rfloor$ is the floor function, defined such that

(3.3) \begin{align} \lfloor x \rfloor = \begin{cases} x, & \text{if } x \in \mathbb{Z} \\ N-1, & \text{if } N-1 \lt x \lt N \text{ and } x \notin \mathbb{Z}, \ N \in \mathbb{Z}. \end{cases} \end{align}

In the surface parametrisation, we also define $n$ as the umbilic factor, which defines the number of sides in the analytical parametrisation, $m$ is an integer that determines the poloidal motion of the sharp umbilic edge and $n_{\textrm {FP}}$ is the field period. The periodicity of the edge is $n_{\textrm {FP}} (m/n)$ – the edge closes on itself after $n/\gcd (n, n_{\textrm {FP}})$ toroidal turns, where $\gcd$ is the greatest common divisor. The parameters $r_1, r_4, r_5$ define the convexity/concavity of a cross-section and how it changes toroidally, $r_4$ determines the magnetic axis torsion and $r_2$ is another parameter that determines the shape of the axis. The shape of such a boundary, the naming convention we use and periodicity of the edge are illustrated in figure 1.

Figure 1.

An umbilic boundary design with $n_{\textrm {FP}}=4, n=3, m=2, \varrho = 0.344$ . All the rest of the parameters $r_j=0$ in (3.1) and (3.2). In (a) we show the full shape and in (b) we present a single field period from the parametrisation. A point $p_i \in \{0, 1, 2\}$ corresponding to the sharp edge at each toroidal cross-section connects to $\mod (p_i+m, 3)$ after a single field period. The sharp edge defined by $R_{\mathrm{US}}(\theta = 0, \zeta ), Z_{\mathrm{US}}(\theta =0, \zeta )$ meets itself after $n = 3$ toroidal turns.

The floor function creates the sharp umbilic edge, while the rest of the parameters grant us enough freedom that we can start with a boundary shape that has good quasisymmetry and allows us to create an umbilic edge with any low-order rational periodicity $n_{\textrm {FP}} (m/n)$ . It is also stellarator-symmetric, i.e. $R_{\mathrm{US}}(\theta , \zeta ) = R_{\mathrm{US}}(-\theta , -\zeta ), Z_{\mathrm{US}}(\theta , \zeta ) = -Z_{\mathrm{US}}(-\theta , -\zeta )$ .

We also define a separate grid on the boundary $\rho = 1$ , the umbilic edge, on which a sharp curvature is imposed, defined by values $(\hat {\theta }, \zeta )$ so that

(3.4) \begin{equation} \hat {\theta } = \frac {m n_{\textrm {FP}} \zeta + \sum _{k}^{N_{\mathrm{u}}} a_{k} \sin (n_{\textrm {FP}} \zeta )}{n}, \quad \gcd (m, n) = 1. \end{equation}

For the surface parametrisation, $\theta = 0$ corresponds to the umbilic edge – initial $\hat {\theta }$ is calculated along this edge but it can change freely during optimisation as the optimiser varies the coefficients $a_k$ . Note that $\theta$ is the DESC poloidal angle throughout this paper, unless stated otherwise.

However, creating and solving with a perfectly sharp edge is difficult with a fully spectral solver. Therefore, in the next section, we explain how we approximate umbilic boundary shapes with the DESC code.

3.2. Approximating umbilic shapes with DESC

DESC uses a Fourier–Zernike spectral representation to solve the force balance equation in the $(\rho , \theta , \zeta )$ coordinate system, where $\rho = \sqrt {\psi /\psi _b},\, \theta = \theta _{\textrm {PEST}} - \varLambda (\rho , \theta , \zeta ), \, \zeta = \phi$ , the cylindrical toroidal angle. The problem is defined in a cylindrical coordinate system $(R, \zeta , Z)$ by decomposing it into Fourier–Zernike spectral bases (see Panici et al. Reference Panici, Conlin, Dudt, Unalmis and Kolemen2023 for more details):

(3.5) \begin{equation} R(\rho ,\theta ,\zeta ) = \sum _{m=-M,n=-N,l=0}^{M,N,L} R_{lmn} \mathcal{Z}_l^m (\rho ,\theta ) \hat {F}^n(\zeta ), \end{equation}

(3.6) \begin{equation} Z(\rho ,\theta ,\zeta ) = \sum _{m=-M,n=-N,l=0}^{M,N,L} Z_{lmn} \mathcal{Z}_l^m (\rho ,\theta ) \hat {F}^n(\zeta ), \end{equation}

where $l,\ m$ and $n$ are the radial, poloidal and toroidal mode numbers, whereas $L,\ M$ and $N$ define the largest values of $l,\ m$ and $n$ , respectively, and $\mathcal{Z}, \hat {F}$ are Zernike and Fourier basis functions. These parameters define the resolution of a DESC equilibrium.

To solve the ideal MHD equation inside the umbilic shape, we have to first fit the Fourier–Zernike basis at $\rho = 1$ to the umbilic boundary, which requires solving a set of linear equations,

(3.7) \begin{eqnarray} \left [\begin{array}{cc} \sum _{m=-M,n=-N,l=0}^{M,N,L} R_{lmn} \mathcal{Z}_l^m (\rho = 1,\theta ) \hat {F}^n(\zeta )\\[4pt] \sum _{m=-M,n=-N,l=0}^{M,N,L} Z_{lmn} \mathcal{Z}_l^m (\rho = 1,\theta ) \hat {F}^n(\zeta ) \end{array}\right ] = \left [\begin{array}{cc} R_{\mathrm{US}}(\theta , \zeta )\\[4pt] Z_{\mathrm{US}}(\theta , \zeta ) \end{array}\right ]\!, \end{eqnarray}

to find $R_{lmn}, Z_{lmn}$ so that the boundary shape matches with the umbilic parametrisation. Similarly, to obtain $a_k$ , we parametrise the umbilic edge using $R_{\mathrm{curve}} = R(\rho =1,$ $\theta =0, \zeta ), Z_{\mathrm{curve}} = Z(\rho = 1, \theta =0, \zeta )$ , and perform an inverse Fourier transform to obtain $a_k$ . Note that upon fitting a spectral basis, the umbilic edge is smoothed out. After representing the umbilic shape in the Fourier–Zernike basis, we solve the ideal MHD equation inside the volume. Figure 2 illustrates this process of fitting and solving force balance in a US131 boundary shape.

Figure 2.

Contour plot of the flux surface average force balance error $|F|$ plotted on the cross-section $\zeta = 0$ of a vacuum DESC equilibrium created using a US131 boundary with $\rho = 0.1,\ r_1=0.133,\ r_5=0.1$ . The actual boundary (in solid black) is obtained after fitting a Fourier–Zernike series to the true umbilic boundary (in dashed blue). The resolution of the DESC equilibrium is $L=M=N=14$ and the resolution of the umbilic curve is $N_{\mathrm{u}} = 12$ .

The sharpness of the umbilic edge increases as we increase the resolution of the fit surface and the agreement between $\texttt {DESC}$ and the umbilic parametrisation improves. Alternatively, one may use a double-Chebyshev (Fortunato Reference Fortunato2024) instead of a double-Fourier representation to improve the spectral fit, local basis methods such as finite element to represent and solve the equilibrium or use a nonlinear transformation to map the circular torus in the Fourier–Zernike basis to an umbilic shape, similar to the transformation used for circular tokamaks by Webster & Gimblett (Reference Webster and Gimblett2009). However, using a new basis and performing equilibrium optimisation are beyond the scope of this work.

In the following section, we discuss how, beginning with the parametrisation of the umbilic surface, we can optimise for omnigenous USs.

3.3. Optimising an umbilic equilibrium with DESC

As we noted in the previous section and figure 2, after fitting a Fourier–Zernike basis to the umbilic boundary, it becomes smooth because approximating a sharp boundary with a double-Fourier representation is usually difficult. Therefore, we have to modify the optimisation process for an umbilic equilibrium to maintain a high curvature, instead of a perfectly sharp umbilic edge. Additionally, for these shapes to be practically feasible fusion reactors, we also need good omnigenity. Therefore, we use the omnigenity and effective ripple objectives to improve the omnigenity throughout the volume.

After fitting to a Fourier–Zernike basis, the high-curvature edge, in general, does not align with the fieldline. In the unoptimised shape defined in (3.1) and (3.2), the $\iota$ is typically too small because of weak shaping. Therefore, we have to optimise separately for a boundary rotational transform to be close $n_{\textrm {FP}} m/n$ . Also, to ensure that the umbilic edge curve matches with the fieldline labelled $\alpha =0$ , we enforce a strong negative curvature $\kappa _{2, \rho }$ along the umbilic edge defined by (3.4). The definitions of the principal curvatures are given in Appendix A. The overall objective function is

(3.8) \begin{eqnarray} \mathcal{F}_{\text{stage-one}} = w_{A}\, f_{\mathrm{aspect}}^2 + w_{\iota }\, f_{\textrm {iota}}^2 + w_{O}\, f_{\mathrm{om}}^2 + w_{U}\, f_{\mathrm{umbilic}}^2 + w_R\, f_{\mathrm{ripple}}^2, \end{eqnarray}

where $f_{x}$ are various objectives on the right-hand side for aspect ratio, boundary rotational transform, omnigenity, high curvature along the umbilic edge and effective ripple and $w_A, w_{\iota }, w_{O}, w_{U}, w_{R}$ are weights used with each objective function. The exact definitions of these objectives are provided in Appendix B. At each iteration of a DESC optimisation, $\mathcal{F}_{\text{stage-one}}$ is minimised while satisfying (2.3). Mathematically,

(3.9) \begin{equation} \min \mathcal{F}_{\text{stage-one}}(\hat {\boldsymbol{p}}), \qquad \textrm {s.t.} \quad \boldsymbol{\nabla }\left (\mu _0 p + \frac {B^2}{2}\right ) - \boldsymbol{B}\boldsymbol{\cdot }\boldsymbol{\nabla }\boldsymbol{B} = 0,\quad \psi _{\mathrm{b}} = \psi _{\mathrm{b}0}, \end{equation}

where $\psi _{\mathrm{b}0}$ is the user-provided enclosed toroidal flux by the boundary and $\hat {\boldsymbol{p}} \in \{R_{\mathrm{b}, mn}, Z_{\mathrm{b}, mn}, a_k\}$ comprises parameters that determine the boundary shape and the position of the umbilic curve on the boundary. With the objective $\mathcal{F}_{\text{stage-one}}$ defined, we run DESC on a single NVIDIA A100 GPU. A single optimisation takes less than $90$ minutes.

4.1. A US131( $n_{\textrm {FP}} = 1, n=3, m =1$ ) omnigenous vacuum stellarator

We begin by calculating an omnigenous vacuum equilibrium with a single field period $n_{\textrm {FP}} = 1$ , $n=3, m=1$ . To do this, we start with the US131 boundary shape, using the parametrisation from § 3.1 with ${\varrho }=0.1\, \textrm {m},\ r_1 = 0.133,\ r_5 = 0.1$ and the rest of the parameters $r_j = 0$ . The toroidal flux enclosed by the initial equilibrium is $\psi _{\textrm {b}} = \pi {\varrho }^2\ (\textrm{Tm}^2)$ . This shape is then approximated by DESC as explained in § 3.2 and optimised as described in § 3.3 with $w_{\textrm{O}} = 0$ Footnote ² to obtain an omnigenous equilibrium with an aspect ratio $A = 10.5$ . The results are presented in figures 3 and 4.

Figure 3.

The US131 optimisation results showing a significant increase in magnetic axis torsion in (a) from the initial (dotted) to the optimised (solid) boundary. This leads to an increase in the rotational transform, as shown in (b) such that $\iota (\rho = 1)$ approaches $1/3$ . The shaping also leads the fieldlines to align with the umbilic edge in the optimised equilibrium as shown in (c).

To ensure that the high-curvature umbilic edge is aligned with the magnetic fieldlines, the boundary undergoes deformation, which results in an increased rotational transform. The optimised design minimises the objective $\mathcal{F}_{\text{stage-one}}$ while ensuring that the normalised second principal curvature is the lowest along the umbilic edge. The curvature on the umbilic edge is in the range $\kappa _{2, \rho } \in [-187, -70]\, {\rm m}^{-1}$ whereas the average curvature on the rest of the boundary is $\mathrm{avg}(\kappa _{2, \rho }) = -12 \,{\rm m}^{-1}$ . This causes the magnetic fieldline to stay close to the umbilic edge throughout the range $\zeta \in [0, 6 \pi ]$ as shown in figure 3(c). In the limit where the edge is perfectly sharp, the fieldline will coincide with the umbilic edge.

The optimised US case is able to achieve an effective ripple $\epsilon _{\mathrm{eff}}$ of less than $2\,\%$ ( $\epsilon _{\mathrm{eff}}^{3/2} \lt 0.0028$ ) for most of the volume $(\rho \lt 0.86)$ and $\epsilon _{\mathrm{eff}}^{3/2} \leqslant 0.001$ for $\rho \leqslant 0.6$ . This ensures that neoclassical transport in the $1/\nu$ low-collisionality regime will not deteriorate the device performance in the core but may deteriorate confinement in the edge. An interesting observation is that we can achieve a low ripple transport despite the lack of a specific helicity of the $B$ contours in the Boozer plot (figure 4 b). This is an instance of a piecewise-omnigenous (Velasco et al. Reference Velasco, Calvo, Escoto, Sánchez, Thienpondt and Parra2024) field that is far from quasisymmetry. Since we did not seek omnigenity with a specific helicity ( $w_{\mathrm{O}} = 0$ ), the effective ripple penalty acted as a proxy for omnigenity, leading to a piecewise-omnigenous configuration. To prove that these configurations are piecewise-omnigenous, we analyse the normalised second adiabatic invariant $\tilde {\mathcal{J}}_{\parallel }$ in Appendix C.

Figure 4.

Output from the vacuum US131 optimisation showing Boozer plots on the boundary and the effective ripple. The output in (b) does not have a specific helicity but still achieves a low ripple transport.

For a vacuum problem, obtaining the right fieldline pitch ( $\iota = n_{\textrm {FP}} m/n$ ) on the boundary and ensuring force balance over a perfectly sharp (unapproximated) boundary make this problem difficult to solve compared with a finite- $\beta$ equilibrium where one can specify the fieldline pitch throughout the volume. Therefore, we have compromised by creating a high-curvature edge instead of a perfectly sharp edge ( $\kappa _{2, \rho } \rightarrow \infty$ ). If one were to use an alternative representation to represent a perfectly sharp boundary and solve this problem exactly, we would observe the rotational transform almost identical to the initial equilibrium in figure 4(c) sharply rising to $\iota = m\, n_{\textrm {FP}}/n$ at the edge.The behaviour of high-shear regions near sharp points would be similar to what is observed in tokamaks.

4.2. Modular coil design for the US131 equilibrium

In this section, we design modular coils for the US131 case using the DESC optimiser and elucidate the various difficulties faced when designing modular coils for strongly shaped USs. We create a coilset of $32$ (stellarator-symmetric with $16$ unique coils) coils using the Python implementation of REGCOIL (Landreman Reference Landreman2017; Panici et al. Reference Panici2025) in DESC. Starting with this initial coilset, we then perform stage-two optimisation, where we fix the equilibrium boundary and allow the shape of the modular coils to change while minimising the objective

(4.1) \begin{eqnarray} \mathcal{F}_{\textrm {stage-two}} = (B_{\textrm {out}}^2 - (B_{\textrm {in}}^2 + 2 \mu _0 p))^2 + (\boldsymbol{B}_{\textrm {out}} \boldsymbol{\cdot }\boldsymbol{\hat {n}})^2 \end{eqnarray}

on the plasma boundary. Note that in stage-two optimisation, we ignore the surface current $\boldsymbol{K} = (\boldsymbol{B}_{\textrm {out}} - \boldsymbol{B}_{\textrm {in}}) \times \boldsymbol{\hat {n}}$ . Therefore, to verify our solution, we calculate the fieldline trajectory due to the coils and create a Poincaré plot. The original fixed-boundary solution, its comparison with the Poincaré cross-section and the optimised coilset are presented in figure 5.

Figure 5.

Coils and $\boldsymbol{B}_{\mathrm{out}}\boldsymbol{\cdot }\hat {\boldsymbol{n}}$ error plotted on the plasma boundary. Here $\mathrm{max}|(\boldsymbol{B}_{\mathrm{out}}\boldsymbol{\cdot }\hat {\boldsymbol{n}})/\boldsymbol{B}_{\mathrm{out}}| \leqslant 1\,\%$ . Only one group of unique coils are plotted. At the upper right we plot the Poincaré cross-section at $\zeta =0$ along with the fixed-boundary solution (in red). The grey curves mark a favourable position to place a divertor. At the lower right we plot the $\boldsymbol{B}_{\mathrm{out}}\boldsymbol{\cdot }\hat {\boldsymbol{n}}$ error on a two-dimensional plane along with the umbilic edge (dashed purple).

Note that even though the curvature along the umbilic ridge is high, it does not behave like a perfect X-point (or X-line) as the fieldline structure becomes chaotic in the edge and resembles the ‘turnstiles’ topology discussed by Punjabi & Boozer (Reference Punjabi and Boozer2022). However, even though the fieldlines do not exit in the regions of the highest curvature, like a tokamak X-point, the Poincaré traces of the fieldlines stay aligned with the plasma boundary and ‘flare’ near the umbilic edge, which still makes the umbilic edge a favourable location to place divertors or limiters. This study indicates that the behaviour of magnetic fieldlines for a high-curvature ridge can differ significantly from that for a perfectly sharp edge.

Another observation is that modular coil design becomes complicated for umbilic equilibria as seen in the figure. The coils assume high curvature to reduce the normal field error along the umbilic edge. Increasing the number of coils improves the agreement of the fixed-boundary solution with the Poincaré cross-section and reduces the maximum coil curvature but also reduces the coil–coil distance. However, the number of modular coils needed for these umbilic equilibria already exceeds the number of coils used for a single field period for a typical stellarator, which is almost always less than 10.

Finally, we note that the regions of maximum normal field error are parallel to the umbilic edge so that, on average, $\boldsymbol{b} \boldsymbol{\cdot }\boldsymbol{\nabla } (\boldsymbol{B}_{\mathrm{out}}\boldsymbol{\cdot }\hat {\boldsymbol{n}}) = 0$ .Footnote ³ This pattern indicates a hidden relationship between the normal field error and the second principal curvature in USs.

To ensure practical coil design for USs, a better technique would be to perform single-stage optimisation, which simultaneously changes the plasma boundary and coil shapes, a union of $\mathcal{F}_{\text{stage-one}}$ and $\mathcal{F}_{\text{stage-two}}$ , accounting for the coil complexity while maintaining omnigenity inside the plasma volume. This is done in § 5.2 for a simpler problem. Another possibility is to use umbilic coils, helically wound divertor-like coils with a fractional helicity ( $m/n$ ) that carry a small current that pushes (pulls) the plasma away from (towards) itself, creating cusp-like regions, leading to a high-curvature umbilic edge. We do a similar optimisation to modify a tokamak in § 5.1, where we add an umbilic coil without changing the shape of the toroidal field coils to create an umbilic edge. In the next section, we use the same optimisation process to create a finite- $\beta$ , omnigenous US.

4.3. A US131 sensitivity analysis of the field structure in the edge

In this section, we introduce a line current source along the magnetic axis of the US131 configuration. The line current aims to simulate an internal fluctuation within the plasma. For different values of the current source strength, we perform the fieldline integration across the entire volume. The purpose of this investigation is to assess the resilience of the fieldline structure at the edge of the US131 stellarator and to determine the extent of modification required for the divertor configuration. We present the fieldline structure for different values of the current in figure 6.

Figure 6.

Poincare plots at $\zeta = 0$ with a current source placed on-axis for different values of the current, compared with the original fixed-boundary solution (in red). Increasing the current significantly reduces the nestedness of the flux surfaces but the fieldline structure in the edge still ‘flares’ out near the high-curvature edges and the divertor placement does not need to be modified.

We observe that even though the flux surfaces inside the plasma boundary lose nestedness and become stochastic, the divertor structure used for the unperturbed US131 case is still an effective choice for this equilibrium, i.e. the fieldline structure in the edge is resilient to fluctuations inside the plasma. Another observation is the emergence of ‘O’-point-like structures near the high-curvature regions in figure 6(b).

4.4. A US252 $(n_{\textrm {FP}}=2, n=5, m=2)$ omnigenous finite- $\beta$ stellarator

In this section, we generate a two-field-period $n=5, m = 2$ finite- $\beta$ stellarator. We start with the umbilic parametrisation defined in § 3.1 with $\varrho =0.1,\ r_1 = 0.133,\ r_5 = 0.1$ and use a pressure profile $p = 5 \times 10^3 (1-\rho ^2)\, \mathrm{Pa}$ , a rotational transform $\iota = 0.755 + 0.043 \rho ^2$ and an enclosed toroidal flux $\psi _{\textrm {b}} = \pi {\varrho }^2 = 0.00314\, \textrm {T m}^2$ . This creates an equilibrium with $\beta = 0.5\,\%$ .

A finite- $\beta$ stellarator provides more freedom compared with a vacuum case because the rotational transform can come from a combination of boundary shaping and plasma current. Therefore, we can achieve a higher rotational transform and use the additional freedom to achieve a much lower ripple transport compared with the vacuum case presented in the previous section. The optimisation uses the same objective function defined in (3.8) with the pressure and iota profiles fixed and $w_{\iota } = 0$ . These results are presented in figure 7.

Figure 7.

Optimisation output showing the (a) initial and (b) final magnetic field strength Boozer plots and (c) the effective ripple. The black lines in (a) correspond to the magnetic fieldlines and the dashed lines correspond to the umbilic curve. The optimised configuration is another instance of piecewise omnigenity.

We can see in figure 7 that the ridges in the Boozer plots are only clearly visible in the initial equilibrium, which is close to quasisymmetry. The magnetic field of the optimised omnigenous equilibrium does not have any visible ridges in the Boozer plot, even though the curvature along the umbilic ridge increases after optimisation. This indicates that the appearance of ridges in the Boozer representation may only be a feature of quasisymmetric stellarators.

This equilibrium is optimised for poloidal omnigenity but, due to the umbilic curvature penalty, limits the poloidal omnigenity to the low-field region. The Boozer plot in figure 7(b) looks qualitatively like a piecewise-omnigenous Velasco et al. (Reference Velasco, Calvo, Escoto, Sánchez, Thienpondt and Parra2024) configuration, which looks similar to the magnetic field in the W7-X stellarator. The piecewise omnigenity of this equilibrium is demonstrated further in Appendix C. We also plot the cross-section of the equilibrium along with the second principal curvature in figure 8. The neoclassical transport coefficient $\epsilon _{\mathrm{eff}}^{3/2} \lt 0.001$ for most ( $\rho \lt 0.7$ ) of the volume, which means the equilibrium is omnigenous in the core.

Figure 8.

Cross-section and three-dimensional shape of the optimised US252 equilibrium. The aspect ratio of the optimised equilibrium is $A = 6.5$ . In (c), we see that around half of the rotational transform of the optimised equilibrium comes from shaping (Hirshman & Hogan Reference Hirshman and Hogan1986) reducing the dependence on external current drive.

We also observe that the optimised equilibrium deforms and rotates as we move toroidally, while maintaining high curvature along the umbilic curve. The curvature along the umbilic edge $\kappa _{2, \rho } \in [-67, -36]\, {\rm m}^{-1}$ while the average curvature on the rest of the plasma boundary is $\mathrm{avg} (\kappa _{2, \rho }) = -8\, {\rm m}^{-1}$ . The variation in the curvature is broad, indicating that imposing high curvature at specific locations may be more suitable than others. This can be readily accomplished by selecting sections of the umbilic curve where high curvature is applied and could be a potential way to place potential locations for non-resonant divertors (Bader et al. Reference Bader, Hegna, Cianciosa and Hartwell2018b ).

5.1. A US111 with an umbilic coil current opposite to the plasma current (counter- $I$ case)

Umbilic coils differ from helical coils by creating a ridge that may serve as a potential location for an X-point. These coils are similar to tokamak divertor coils, albeit without axisymmetry. To demonstrate their applicability, we modify an $n=1$ MHD-stable equilibrium from the HBT-EP experiment ( $R_0 = 0.94\, {\rm m},\ A = 7.5$ ) by adding an $m = 1, n = 1$ ridge to it, breaking its axisymmetry, and then design an umbilic coil for it without modifying the toroidal or vertical field coils. The specific details of the equilibrium used are given in Appendix D.

To accomplish this, we start by using $20$ toroidal field (TF)Footnote ⁴ coils and $4$ vertical field poloidal (PF) coils as done in the original HBT-EP experiment (Gates Reference Gates1993; Sankar et al. Reference Sankar1993) and add an $m = 1, n =1$ umbilic coil outside the plasma boundary at a distance of $1.5$ times the minor radius. Similar to § 4.2, we perform second-stage optimisation to find the right shape of the umbilic coil. To ensure that the coils from stage-two optimisation give expected results, we calculate the free-boundary solution and present it in figure 9.

Figure 9.

Output from the stage-two optimisation and free-boundary solution. (a) The full coilset of the modified HBT-EP experiment with the magnetic field strength on the boundary. (b) The umbilic coil and the curvature $\kappa _{2, \rho }$ on the boundary. (c) The cross-section of the LCFS from the free-boundary DESC solver along with the umbilic coil at four different toroidal angles and the directions in which they carry current relative to the plasma.

The effect of the umbilic coil can be decomposed as a rigid shift of the whole equilibrium and a small deformation of the boundary. Since the rigid shift (Sengupta et al. Reference Sengupta, Nikulsin, Gaur and Bhattacharjee2024) and the size of the ridge are small compared with the minor radius, it does not compromise omnigenity (as seen later in figure 12 a). For a perfectly sharp umbilic edge, the magnetic fieldline must coincide with the edge, but even when the ridge lacks perfect sharpness, the magnetic fieldline stays close to it due to its high curvature. The curvature along the umbilic ridge $\kappa _{2, \rho } \in [-15, -32]\, {\rm m}^{-1}$ whereas the average curvature $\mathrm{avg}(\kappa _{2, \rho }) = -7\, {\rm m}^{-1}$ .

The TF coils each carry a current of $105\,\text{kA}$ , the PF coils carry $29.8\,\text{kA}$ and the umbilic coil has a current $I_{\mathrm{umbilic}} = 2.1\,\text{kA}$ . But due to the proximity of the umbilic coil compared with TF and PF coils, it can still affect the plasma near the boundary significantly. As we can see, the umbilic coil is able to create an $m=1, n=1$ high-curvature ridge. The ridge formation is entirely attributed to the umbilic coil, as all other coils are axisymmetric. The current flowing through the coil is in the opposite direction to the plasma current. However, due to the three-dimensional nature of this set-up, the plasma still bulges towards the coil.

To understand the structure of the magnetic field near the edge, we use the FIELDLINES (Lazerson et al. Reference Lazerson2016) code from the STELLOPT (Lazerson et al. Reference Lazerson, Schmitt, Zhu and Breslau2020) optimiser. We observe the presence of a stochastic layer outside the plasma boundary, similar to heliotron/torsatron configurations. Since the plasma and the umbilic coil are carrying currents in opposite directions, the poloidal field cannot vanish between them, which is why we do not observe an X-point. Therefore, the umbilic coil can create and vary the curvature of the ridge without creating and varying the current in the umbilic coil without creating an X-point. This also affects the fraction of the rotational transform generated by the umbilic coils. These details are provided in figure 11.

Figure 10.

Poincaré section at different toroidal angles compared with the free-boundary DESC solution (red). The red cross marks the position of the umbilic coil.

Figure 11.

Free-boundary solution at $\zeta = 0$ for (a,b) different values of the umbilic coil current $I_{\mathrm{umbilic}}$ and (c) the fraction of the rotational transform generated by the umbilic coil current.Footnote ⁵

By controlling the current in the umbilic coil, we can shape the plasma without moving the limiter. Also, the ability to vary $\iota$ would help to stabilise various MHD instabilities and the $m=1, n=1$ tearing mode as done in the RFX-mod experiment (Piron et al. Reference Piron2016) or experiments like the Madison Spherical Torus, where the edge rotational transform close to one can be achieved (Hurst et al. Reference Hurst2022). The J-TEXT (Liang et al. Reference Liang2019) tokamak has recently proposed a similar idea (Li et al. Reference Li2024) of using umbilic coils to stabilise the edge of a tokamak.

Upon breaking the axisymmetry of the magnetic field, one must be careful as increasing the current $I_{\mathrm{umbilic}}$ for the same coil shape would degrade omnigenity and hence neoclassical transport. To understand the effect of umbilic coils on omnigenity, we present Boozer plots for the three values of current in figure 12.

Figure 12.

Magnetic field strength in Boozer coordinates on the boundary $(\rho =1.0)$ for different values of the umbilic coil current and the fieldline $\alpha = 0$ . As the curvature of the ridge increases, so does the localised distortion of the $|\boldsymbol{B}|$ contours close to the fieldline.

We can see that the omnigenity degrades as we increase $I_{\mathrm{umbilic}}$ but the neoclassical losses are still acceptable: $\epsilon _{\mathrm{eff}}^{3/2} \lt 0.001$ . To control instabilities and transport due to MHD modes, we may need to generate a higher $\iota _{\mathrm{shaping}}$ while ensuring omnigenity. One way to explore multiple such coil and equilibrium configurations would be to perform a single-stage optimisation, which simultaneously changes the plasma boundary and coil shapes. This is beyond the scope of the current study.

5.2. A US111 with an umbilic coil current in the same direction as the plasma current (co- $I$ case)

In a manner similar to that of the previous section, we create a set-up where we modify the HBT-EP experiment and add an $m = n = 1$ ridge to the boundary. However, we do this while ensuring that the toroidal component of the current in the umbilic coil $I_{\mathrm{umbilic}}$ is in the same direction as the plasma current $I = -35.2\, {\rm kA}$ . To find the umbilic coil shape, coil current and the corresponding plasma shape that minimises the normal field error $\boldsymbol{B}\boldsymbol{\cdot }\hat {\boldsymbol{n}}$ and the total pressure across the plasma boundary, we combine the objectives defined in (3.8) and (4.1) in a process known as single-stage optimisation. We solve

(5.1) \begin{align} &\min (\mathcal{F}_{\text{stage-one}} + \mathcal{F}_{\text{stage-two}}), \quad \textrm {s.t.} \ \boldsymbol{\nabla }\left (\mu _0 p + \frac {B^2}{2}\right ) - \boldsymbol{B}\boldsymbol{\cdot }\boldsymbol{\nabla }\boldsymbol{B} = 0,\quad \psi _{\mathrm{b}} = \psi _{\mathrm{b}0},\nonumber \\ & p = p_{0}(\psi ),\quad I = I_0(\psi ), \end{align}

where $\psi _{\mathrm{b}0}$ is the user-provided enclosed toroidal flux by the boundary and $p_0(\psi )$ and $I_0(\psi )$ are the pressure and toroidal current profiles. The optimisable set of parameters is $\hat {\boldsymbol{p}} \in \{R_{\mathrm{b}, mn}, Z_{\mathrm{b}, mn}, a_k, X_n, Y_n, Z_n, I_{\mathrm{umbilic}}\}$ , where $X_n, Y_n, Z_n$ are the Fourier coefficients that determine the shape of the umbilic coil. The coilset and free-boundary solution are presented in figure 13. After optimisation, the normalised $\mathrm{max}(\boldsymbol{B}\boldsymbol{\cdot }\hat {\boldsymbol{n}}) \lt 0.5\,\%$ and the total pressure difference is $\lt 0.1\,\%$ . Because we allow the boundary shape to change, we have less control over the curvature along the umbilic edge, so there is a lot of variation. The curvature along the edge lies in the range $\kappa _{2, \rho } \in [-25, -13]\, {\rm m}^{-1}$ with the average curvature $\mathrm{avg}(\kappa _{2, \rho }) = -8\, {\rm m}^{-1}$ . The neoclassical transport coefficient ${\epsilon _{\mathrm{eff}}}^{3/2} \lt 0.001$ implies that the magnetic field is still omnigenous. Finally, we calculate the rotational transform profile of the optimised equilibrium to quantify the effect of the umbilic coil in figure 14.

Figure 13.

Coilset and free-boundary solution for the case where the current in the umbilic coil is in the same direction as the plasma current. The current carried by the umbilic coil is $I_{\mathrm{umbilic}} = -3.1\,{\rm kA}$ . All coils but the umbilic coil are axisymmetric.

Figure 14.

The total rotational transform profiles and the contribution from shaping and plasma current. The non-axisymmetric shaping is purely a consequence of the umbilic coil which contributes to approximately $5\,\%$ of the total rotational transform in the edge.

The effect on the rotational transform profile due to the umbilic coil rises sharply as we move towards the edge. Therefore, it is possible that outside the plasma boundary the umbilic coil is able to affect the total magnetic field in such a way that it forms an X-point.